Accretion discs

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Accretion discs are flattened astronomical objects made of rapidly rotating gas which slowly spirals onto a central gravitating body. The gravitational energy of infalling matter extracted in accretion discs powers stellar binaries, active galactic nuclei, proto-planetary discs and some gamma-ray bursts. The black hole accretion in quasars is the most powerful and efficient stationary engine known in the universe. In accretion discs the high angular momentum of rotating matter is gradually transported outwards by stresses (related to turbulence, viscosity, shear and magnetic fields). This gradual loss of angular momentum allows matter to progressively move inwards, towards the centre of gravity. The gravitational energy of the gaseous matter is thereby converted to heat. A fraction of the heat is converted into radiation, which partially escapes and cools down the accretion disc. Accretion disc physics is thus governed by a non-linear combination of many processes, including gravity, hydrodynamics, viscosity, radiation and magnetic fields.

The observable physical quantity of radiation produced in accretion discs is the luminosity. As photons carry momentum and thus can exert pressure there is a maximum possible luminosity at which gravity is able to balance the outward pressure of radiation. The limit for a steady, spherically symmetric accretion flow is given by the Eddington luminosity,

\( L_{\rm Edd} = G M \frac{4 \pi m_H \, c}{\sigma_T} = 1.26 \times 10^{38} \; (M/M_{\odot}) \; {\rm erg/s}, \)

where G is the gravitational constant, c is the speed of light, M is the mass of the gravitating body, \(m_H\) and \(M_{\odot}\) are the hydrogen and solar mass, respectively, and \(\sigma_T\) is the Thompson cross-section. The Eddington limit is used as a unit to quantify the luminosity of an object. Since accretion discs are not spherical and often have additional stresses (see above) that can counteract the radiation pressure along with gravity, they may be brighter than this limit and radiate at super-Eddington luminosity.

A short summary of the basic properties of accretion discs
    based on a lecture by Kristen Menou   (November 2008, Nordita, Stockholm, Sweden)


white dwarfs (WD)
in cataclysmic binaries

black hole (BH) or
neutron star (NS)
quasars and
other AGN
gamma ray burst
(GRB) sources
Protoplanetary tw hya 2013.jpg Wz-sge-disc-spirals.jpg Accretion-disc.jpg M87-jet.jpg Gamma-ray.jpg

Basic physics

The central part of a dense molecular cloud collapses to a proto-star surrounded by a proto-planetary accretion disc. Self gravity and sedimentation trigger the formation of planets. Bipolar outflows (slow jets) often emerge from proto-planetary discs.
U Gem is the prototype of a dwarf novae system, i.e. a close stellar binary, with "primary" being a WD with accretion disc. The disc's brightness in the visible light increases 100-fold every ~120 days and returns to the original level after a ~week, due to (mainly) a limit-cycle instability.
X-ray binaries (XRB) consist a mass loosing main-sequence "secondary" star and accreting BH or NS. Among XRBs, the soft X-ray transients (with BH or NS) show quasi-periodic outbursts. Most of the BH XRBs exhibit fast jets, and for this reason are called microquasars.
AGN are supermassive BH at centres of galaxies. Accretion produces radiative power that often outshines the host galaxy. The accretion disc is surrounded by moving gas clouds and encircled by a large torus of gas and dust. Very fast (almost speed of light) jets emerge from many AGNs.
GRBs are the most energetic explosions in the universe. Models of GRBs invoke a black hole (\(M \sim 3 - 10 M_{\odot}\)) accreting matter at highly super-Eddington rates. The huge power of gamma-rays is possibly due to an extraction of the BH rotational energy (the Blandford- Znajek mechanism).

Accreting central object

\(M_{star} = 1 M_{\odot}\)
\(M_{WD} = 1 M_{\odot}\)
\(M_{BH} = 3-20 M_{\odot} \)
\( M_{NS} = 1-2 M_{\odot}\)
\(M_{BH} = 10^{6} - 10^{9} M_{\odot}\)
\(M_{BH} = 3-20 M_{\odot}\)

Disc size

disc = \(10^{11} - 10^{15} \) cm

disc = \(10^{9} - 10^{10} \) cm

disc = \(10^{6} - 10^{11}\) cm

disc = \(10^{6} - 10^{11}\) cm [\(M/M_{\odot}\)]

disc = \(10^{5} - 10^{?}\) cm [\(M/M_{\odot}\)]

Midplane temperature (inner - outer disc)

\(T_c = 10^3 - 10^1\) K \(T_ = 10^5 - 10^3\) K \(T_c = 10^7 - 10^3\) K \(T_c = 10^5 - 10^2\) K \(T_c = 10^{10} - 10^9\) K


\(L << L_{\rm Edd}\) \(L << L_{\rm Edd}\) \(L << L_{\rm Edd}\)
\( L ~ L_{\rm Edd}\)
\(L < L_{\rm Edd}\)
\(L > L_{\rm Edd}\),   \(L >> L_{\rm Edd}\)
\(L >> L_{\rm Edd}\)

Theoretical disc models

Mostly thin discs, thick discs (early epochs), layered discs (with a magnetically inactive 'dead zone' in the mid-plane region)
Thin discs (truncated and with funnel/column accretion if the WD is magnetised)
Thin discs, slim discs, ADAF
Thick discs (corona), slim discs
Thick discs, thin discs, hyper-accretion, NDAF

Angular momentum transport

Radial: In the inner disc region and at the surface, where the disc is sufficiently ionised (by X-rays, cosmic rays and collisions), via MRI induced turbulence; in the dead zone via gravitational instability
Vertical: Via outflows and/or torque exerted by large scale magnetic fields.
Local: In the high state via MRI induced turbulent viscosity;
Global: Direct dissipation by tidal spirals when the incoming supersonic flow shocks on the accretion disc
Local : MRI drives a turbulent viscosity which also produces shear stresses;
Global: Spiral shocks (?)
Inner disc: Viscous friction (MRI);
Outer disc: Possibly by global disturbances in the gravitational field (gravito-turbulence)
Inner disc: MRI induced turbulent viscosity (in the optically thick mid-plane a very large neutrino viscosity could shut off MRI);
Outer disc: (\( > 140 R_G\)) uncertain

Cooling processes

Blackbody radiation, convection, collisions
Blackbody radiation
Thin/Slim discs: Advection, blackbody radiation;
ADAF: Advection, bremsstrahlung, Compton scattering
Thick discs (corona): Bremsstrahlung, Compton scattering;
Thin discs: Blackbody radiation
Inner disc: Neutrino emission;
Outer disc: Advection


Hartmann (1998)
Alexander (2008)
Frank, King & Raine (2002)
Warner (2003)
Frank, King & Raine (2002)
Remillard & McClintock (2006)
Krolik (1998)
on-line compilation
Popham, Woosley & Fryer (1999)
Di Matteo, Perna & Narayan (2002)

Accretion discs in the Universe

Accretion discs in young stellar objects (YSOs)

Figure 1: Accretion disc and jet in a proto-star HH30 observed by the Hubble Space Telescope: the jet (in red) is perpendicular to the accretion disc, seen edge-on (a dark region between two bright lobes), © Burrows, STSci/ESA, WFPC2, NASA)
*K. Wood, M.J. Wolff, J.E. Bjorkman, B. Whitney The HH30 Spectrum: Constraining Circumstellar Dust
Figure 2: A proto-star star in NGC 1333. Reconstruction of a possible look of a proto-planetary disc based on a Spitzer Space Telescope image. An evidence was found for water vapor in the surrounding area, which appears to be one of the key moments in the development of a planetary system around such a star: icy material is falling from the envelope that birthed the star onto a dense, surrounding disc. Credit: NASA/JPL-Caltech/R. Gutermuth (Harvard-Smithsonian Center for Astrophysics)

During star formation, the central part of a dense molecular cloud collapses to a proto-star with a gaseous envelope that finally settles to a rotating proto-planetary accretion disc. Sedimentation and self-gravity in such discs trigger the formation of planets and planetary systems. Proto-stars are heavily embedded in surrounding gas and dust and for this reason visible only in the infrared, millimetre or sub-millimetre wavelength bands. Most of the material that goes into forming a star is accreted through a circumstellar disc and in this process the proto-stellar system drives an energetic bipolar jet and outflow into its surroundings. The least evolved proto-stars are surrounded by remnant proto-planetary accretion discs.

Based on the spectral energy distribution in the infrared and visible light, YSOs are divided into five classes (0-IV), associated with their evolutionary stages. Class 0 refers to collapsing molecular clouds, proto-planetary discs exist in classes I-III, and class IV contains the zero-age main-sequence star.

A related issue: the extrasolar planets

Accretion discs in cataclysmic variables (CVs)

Figure 3: U Gem system, as it would be seen from the Earth. In reality, the image of the system is unresolved, and only the total flux from the secondary (red) and accretion disc (blue) is observed. The primary white dwarf is located at the centre of the accretion disc (too small to be seen). The secondary and the accretion disc periodically eclipse each other, which results in periodic variations in the observed flux (photometry) and spectral features (spectroscopy). From these variations Smak (1971) reconstructed size and shape of the accretion disc.

CVs are binary star systems consisting of a white dwarf ("primary") and a normal star ("secondary", or "companion"). Typically, the CVs have sizes comparable to the Earth-Moon system, and orbital periods of a few hours. When the outer layers of the companion overflow the "Roche lobe", the companion loses matter through the first Lagrangian point \(L_1\) of the rotating binary system. When the white dwarf is only weakly magnetised, the matter forms an accretion disc around it and eventually reaches its surface. "Dwarf novae" (DN) are CVs that show outbursts lasting for about a week and separated by weeks to months of quiescence. U Gem is the prototype of dwarf novae. The brightness in the visible light of U Gem increases 100-fold every 120 days or so, and returns to the original level after a week or two. The DN phenomenon is due to a specific accretion disc limit-cycle instability, tidal torques, and fluctuations in the mass-transfer rate from the secondary. The geometry of accretion is very different in magnetic CVs, where accretion discs are either truncated or absent and accretion occurs along the magnetic field lines.

There is a solid observational evidence for accretion discs in CVs based on very accurate photometry and spectroscopy.

Accretion discs in Quasars and other active galactic nuclei (AGN)

Most galaxies have supermassive (millions to billions of solar masses) black holes at their centres (nuclei). In AGN, the black hole accretion produces radiative power that usually outshines its host galaxy. The accretion disc is surrounded by a hot corona which contains clouds of gas. Fast moving clouds close to the disc produce broad lines and slow moving clouds further away from the disc produce narrow lines in the AGN spectra. The observational appearance of an AGN may be affected by the presence of a large outer dust torus.

Figure 4: AGN unification scheme. Green arrows indicate the AGN type that is seen from a certain viewing angle. Image credit: NASA
Property Quasars Seyfert Galaxies

Radio Galaxies

Galaxy type Spiral, Elliptical Spiral Giant Elliptical Elliptical
Appearance compact, blue compact, bright nucleus elliptical bright, star-like
Maximum luminosity 100-1,000 Milky Way comparable to bright Spirals strong radio 10,000 Milky Way
Continuum spectrum non-thermal non-thermal non-thermal non-thermal
Absorbtion lines yes none yes none
Variability days to weeks days to weeks days hours
Radio emission some weak strong weak
Redshifts z > 0.5 z ~ 0.5 z < 0.05 z ~ 0.1

AGN are generally divided into two families, the "radio loud" and the "radio quiet", depending on whether they exhibit jets or not. In each family several types of AGN are distinguished by their emission properties. Among them, the spiral galaxies with broad and narrow emission lines (Seyfert 1) or galaxies with just narrow emission lines (Seyfert 2) and their counterparts in the radio loud family, the broad line and/or narrow line radio galaxies. The absence of broad lines in type 2 galaxies has been attributed to a partially obscured inner disc by the outer dust torus when the viewing angle is above ~60°. Since high luminosity Seyfert galaxies are preferably of type 1 the alternative scenario of a receding/advancing torus being responsible for type 1/type 2 is plausible (C. Simpson, 1998). The most luminous beacons in the universe are the radio loud and radio quiet quasars (i.e. quasi-stellar objects). They are observed up to highest redshifts, implying cosmological distance and gigantic energy output. Therefore, despite their name, quasi-stellar objects are anything but stars. However, because quasars shine at such large distances, it is not possible to resolve the bright core. Recent observations detect jets and nebulosity around some of them.

Accretion discs in Microquasars and X-ray binaries

Figure 5: Microquasars found in several X-ray binaries in our Galaxy are scaled down versions of quasars, as pointed out by Felix Mirabel, who also coined the name "microquasar". This figure first appeared in several of Mirabel's articles.
Figure 6: Black hole X-ray binaries in our Galaxy. The figure shows the companion star and the accretion disc drawn to scale. For a comparison, the Sun - Mercury distance is shown. (Figure by Jerome Arthur Orosz).

Main-sequence stars in orbit around an accreting neutron star or black hole are called neutron star binaries (NSB) or black hole binaries (BHB), respectively, and are common objects in a galaxy. Young neutron stars are often strongly magnetised so that their accretion discs are truncated by the magnetic field or do not exist at all. In these cases matter is lead by partial or total column accretion to the neutron star.

Neutron stars and black holes are more compact than white dwarfs. NSB and BHB show thus, due to their stronger gravitational potential, a much higher energy output in comparison to CVs: Their spectral energy distribution is observed up to the X-ray regime. For this reason they are also called X-ray binaries (XRB). Depending on the mass of the companion star, XRB are roughly divided into two categories:

  • Low-mass X-ray binaries, LMXB (in Figure 6, systems with coloured companion stars)
  • High-mass X-ray binaries, HMXB (in Figure 6, systems with white companion stars)

The HMXB family has two sub-classes, (i) the soft X-ray transients that can contain a NS or a BH and exhibit quasi-periodic outbursts and (ii) the pulsars that contain a NS and emit a bean of radiation along the axis of the magnetic field (not necessarily aligned with the NS rotational axis).

Many, if not all, black hole X-ray binaries go through periods during which they emit relativistic twin jets that propagate along the rotational axis of the compact object. They then look like scaled down quasars and are thus often called microquasars.

Accretion discs in gamma ray bursts (GRBs)

Figure 7: Afterglow of the gamma-ray burst captured in X-ray (left) and UV/optical (right) on 19.03.2008. At peak, this GRB was visible with the naked eye. (Credit: NASA/Swift/Stefan Immler)

The most energetic explosions seen in the universe are gamma-ray bursts. They are short, collimated flares of low-energy \(\gamma\)-rays with relativistic emission simultaneously also at longer wavelengths (e.g., X-ray flashes, radio jets). Observations indicate that GRBs are cosmological and followed by slowly fading afterglows. The duration of GRB prompt emission can last from 0.01 - 2 seconds (short bursts) up to 2 - 500 seconds (long bursts) and may be explained by merging compact objects or failed supernovae (collapsars), respectively. Afterglows on the other hand are observed and monitored from a couple of days up to several years. All evidence on the origin of the inner engines (i.e., mergers, collapses, pulsars) of GRBs is deduced indirectly. Energetic requirements suggest, however, a similar configuration of the end products: the formation of a solar-mass black hole surrounded by a massive debris disc (~ 0.1 Msun) with a huge accretion rate. The time scale of the burst is determined by the accretion time of this disc. According to these time scales accretion discs in GRBs are most likely hyper-accreting. This means, the temperatures and densities at the required accretion rates are such, that neutrino production is switched on and the electrons are mildly relativistic and degenerate. Generally, GRBs seem to show similarities to radio-loud AGN and galactic microquasars, since all these systems eject strongly collimated, more or less relativistic flows of matter and involve accretion onto a black hole.

Basic physics of accretion discs

The black hole gravity

The black hole gravitational field is described by three parameters: mass \(M\ ,\) angular momentum \(J\) and charge \(Q\ .\) It is convincingly argued that the astrophysical black holes relevant for accretion discs are uncharged, \(Q = 0\ .\) They are described by the stationary and axially symmetric Kerr geometry, with the metric \(g_{\mu\nu}\) given in the spherical Boyer-Lindquist coordinates \(t, \phi, r, \theta\) by the explicitly known functions of the radius \(r\) and the polar angle \(\theta\ ,\) which are listed in the table below. The table also gives the contravariant form of the metric, \(g^{\mu\nu}\ ,\) defined by \(g^{\mu\beta}\,g_{\nu\beta} = \delta^{\mu}_{~\nu}\ .\) It is defined, \(\Delta = r^2 - 2Mr + a^2\ ,\) \(\Sigma = r^2 + a^2\cos^2\theta\ .\) The signature \((+\,-\,-\,-)\) is used.

The mass and angular momentum have been rescaled into the \(c = G = 1\) units, \(M \rightarrow GM/c^2\ ,\) \(J \rightarrow a = J/c\ .\) For a proper black hole solution it must be \(\vert a \vert \le M\ ,\) and the metric with \(\vert a \vert > M\) corresponds to a naked singularity. The Penrose cosmic censor hypothesis (unproved) states that there are no naked singularities in the Universe.


  \(t\) \(\phi\) \(r\) \(\theta\) \(t\) \(\phi\) \(r\) \(\theta\)
\(t\) \(1 - 2\,M\,r/\Sigma\) \(4\,M\,a\,r\sin^2\theta/\Sigma\) \(0\) \(0\)

\((r^2 + a^2)^2/\Sigma\,\Delta\)
\(- a^2\Delta\sin^2\theta/\Sigma\,\Delta\)
\(2M\,\,a\,r/\Sigma\,\Delta\) \(0\) \(0\)
\(\phi\) \(4\,M\,a\,r\sin^2\theta/\Sigma\)

\(-(r^2 + a^2)\sin^2\theta\)

\(0\) \(0\)



\(0\) \(0\)
\(r\) \(0\) \(0\) \(-\Sigma/\Delta\) \(0\)

\(0\) \(0\) \(-\Delta/\Sigma\) \(0\)
\(\theta\) \(0\) \(0\) \(0\) \(-\Sigma\)

\(0\) \(0\) \(0\) \(-1/\Sigma\)
In any stationary and axially symmetric spacetime, and in particular in the Kerr geometry, for matter rotating on circular orbits with four velocity \(u^{\nu} = (u^t, u^{\phi})\) it is \(\Omega = u^{\phi}/u^t\) and \(j = - u_{\phi}/u_t\ ,\) from which (and \(u^{\nu}\,u_{\nu} = 1 \)) it follows that,
\[ \Omega = -\frac{j\,g_{tt} + g_{t\phi}}{j\,g_{t\phi} + g_{\phi\phi}}, ~~~ j =-\frac{\Omega\,g_{\phi\phi} + g_{t\phi}}{\Omega\,g_{t\phi} + g_{tt}}, ~~~ U_{\rm{eff}} = -\frac{1}{2} \ln \left( g^{tt} - 2j\,g^{t\phi} + j^2\,g^{\phi \phi}\right). \] \((3.2)\)
From equations (3.1), (3.2) and \(dR^2 = g_{rr}dr^2\ ,\) \(dZ^2 = g_{\theta\theta}d\theta^2\ ,\) one derives that the Keplerian frequency \(\Omega_K\) and the two epicyclic frequencies (radial \(\omega_R\) and vertical \(\omega_Z\)) equal
\[ \Omega_K = \frac{c^3}{GM}\left( {r_*}^{3/2} + {a_*}\right)^{-1}, ~~~ \omega_R^2 = \Omega_K^2 \left( 1 - 6{r_*}^{-1} + 8{a_*}\,{r_*}^{-3/2} - 3{a_*}^2\,{r_*}^2 \right), ~~~ \omega_Z^2 = \Omega_K^2 \left( 1 - 4{a_*}\,{r_*}^{-3/2} + 3{_*a}^2\,{r_*}^{-2} \right). \] \((3.3)\)
Here the dimensionless \({r_*}\) and \(\vert{a_*}\vert \le 1\) are defined by \({r_*} = rc^2/GM\ ,\) \({a_*} = Jc/GM^2\ .\) In the strong gravity, i.e. for \({r_*} \sim 1\ ,\) the three frequencies scale as \(1/M\ .\) The radial epicyclic oscillations, \(\delta r(t) \sim \exp(- i\,\omega_R\,t)\ ,\) become dynamically unstable at \(r < \rm{ISCO}\), because there \(\omega_R^2(r) < 0\ .\)

Stable circular Keplerian orbits exist only with radii greater than the radius of ISCO (the innermost stable circular orbit radius). All Keplerian orbits closer to the black hole than ISCO are unstable: without an extra support by non-gravitational forces (i.e. pressure or magnetic field) matter cannot stay there orbiting freely, but instead it must fall down into the black hole. This strong-field property of Einstein's gravity, absent in Newton's theory, is the most important physical effect in the black hole accretion disc physics.

Dynamical, thermal and viscous processes

The accretion discs physics is governed by a non-linear combination of many processes, including gravity, hydrodynamics, viscosity, radiation and magnetic fields. According to a semi-analytic understanding of these processes developed over the past thirty years, the high angular momentum of matter is gradually removed by viscous stresses and transported outwards. This allows matter in the accretion disc to gradually spiral down towards the gravity centre, with its gravitational energy degraded to heat. A fraction of the heat converts into radiation, which partially escapes and cools the accretion disc. Accretion disc physics is often described in terms of dynamical, thermal and viscous processes that occur at different timescales \(t_{\rm dyn}\ ,\) \(t_{\rm the}\ ,\) \(t_{\rm vis}\ :\)
  • Dynamical processes occur with the timescale \(t_{\rm dyn}\) (a time in which pressure force adjusts to combined gravitational and centrifugal forces).
  • Thermal processes occur with the timescale \(t_{\rm the}\) (a time in which the entropy redistribution occurs due to dissipative heating and cooling processes (in particular radiation).
  • Viscous processes occur with the timescale \(t_{\rm vis}\) (a time in which angular momentum distribution changes due to torque caused by dissipative stresses).
In most analytic models it is assumed that \(t_{\rm dyn} < t_{\rm the} < t_{\rm vis}\ ,\) and in the thin disc analytic models it is assumed that \(t_{\rm dyn} \ll t_{\rm the} \ll t_{\rm vis}\ .\) Although neither these inequalities, nor even the very existence of the separate timescales \(t_{\rm dyn}\ ,\) \(t_{\rm the}\ ,\) \(t_{\rm vis}\ ,\) could be considered as well established facts (indeed some of the supercomputer simulations of accretion seem to challenge that), the present understanding of the accretion disc physics --- both in general and in details --- is based (explicitly or implicitly) on this separation into dynamical, thermal and viscous processes.

Dynamical processes, with the timescale \(t_{\rm dyn}\)

Dynamical equilibria of accretion flows are governed by the balance of four forces: gravitational \(\bar G\ ,\) centrifugal \(\bar C\ ,\) pressure \(\bar P\ ,\) and magnetic \(\bar M\ .\) In particular, accretion discs, are characterised by a significant contribution of \(\bar C\ .\) Thus, in accretion discs the angular momentum of matter is high (and therefore dynamically important \(\bar C \sim \bar G + \bar P + \bar M\)) in contrast to another important type of accretion flows --- the quasi-spherical "Bondi" accretion, where the angular momentum is everywhere smaller than the Keplerian (and therefore dynamically unimportant, \(\bar C \ll \bar G + \bar P + \bar M\)). Some authors take this difference as a defining condition: in an "accretion disc" there must be an extended region where the matter's angular momentum is not smaller than the Keplerian angular momentum in the same region. This is illustrated in Figure 8.

"Keplerian" refers to the angular momentum of a fictitious free particle placed on a free circular orbit around the accreting object. According to Newton's theory (applicable to weak gravity), the Keplerian angular momentum at a distance \(r\) from a spherical object with the mass \(M\) equals \((GMR)^{1/2}\ ,\) i.e. it is monotonically increasing, indicating (Rayleigh's) stability of all orbits. According to Einstein's theory, in the strong gravity near a compact object such as a black hole or a neutron star, the Keplerian angular momentum has a minimum at the radius \(r = r_{\rm{ISCO}}\) (see Figure 1). All orbits with \(r > r_{\rm ISCO}\) are stable, all orbits with \(r < r_{\rm ISCO}\) are unstable, the orbit at \(r = r_{\rm ISCO}\) is called the innermost stable circular orbit (ISCO). Even closer to the black hole, for \(r < r_{\rm MB}\ ,\) the unstable orbits are also unbound. For a non-rotating black hole

\[r_{\rm ISCO} = 6GM/c^2\] and \[r_{\rm MB} = 4GM/c^2\ .\]

The existence of ISCO makes physics of the inner part of accretion discs in strong gravity fundamentally different from physics of accretion in weak gravity.
Figure 8: The "Bondi-like" and "disc-like" accretion flows
Most of the accretion disc types (except proto-planetary and GRB ones) have a negligible self-gravity: the external gravity of the central accreting object dominates. The external gravity is important in shaping several crucial aspects of the internal physics of accretion discs, including their characteristic frequencies (that are connected to several important timescales) and their size (inner and outer radius). The most fundamental gravity's characteristic frequencies are the Keplerian orbital frequency \(\Omega_{\rm K}\ ,\) the radial epicyclic frequency \(\omega_r\ ,\) and the vertical epicyclic frequency \(\omega_z\ .\) They are directly relevant for motion of free particles and also play a role for determining equilibria and stability of rotating fluids. In both Newton's and Einstein's gravity the three frequencies are derived from the effective potential \(U_{\rm eff}(r, j)\ ,\) and given by the same formulae,

\[ \left[ \left(\frac{\partial U_{\rm eff}}{\partial r}\right)_j = 0 \right] \rightarrow \left[ \Omega_{\rm{K}}^2 = \Omega_{\rm{K}}^2(r) \right], ~~~\omega_r^2(r) = \left(\frac{\partial^2 U_{\rm eff}}{\partial r^2}\right)_j, ~~~\omega_z^2(r) = \left(\frac{\partial^2 U_{\rm eff}}{\partial z^2}\right)_j, \]


where \(j\) is the specific angular momentum, and derivatives are taken at the symmetry plane \(z = 0\ .\) Small (epicyclic) oscillations around the circular orbit \(r = r_0 = const\ ,\) \(z = 0\) are governed by \( \delta{\ddot r} + \omega^2_r\,\delta r = 0\ , \) \( \delta{\ddot z} + \omega^2_z\,\delta z = 0\ , \) with solutions \( \delta{r} \sim \exp( -i\omega_r t)\ , \) \( \delta{z} \sim \exp( -i\omega_z t)\ , \) which are unstable when \(\omega^2_r < 0\) or \(\omega^2_z < 0\ .\) In Newton's gravity \(U_{\rm eff} = \Phi + j^2/2r\ .\) A spherical Newtonian body has the gravitational potential \(\Phi = -GM/r\ .\) Thus, in this case, \(\Omega_{\rm K}^2 = \omega_r^2 = \omega_z^2 = GM/r^3 > 0\ ,\) i.e. all slightly non-circular orbits are closed and all circular orbits are stable.

In Einstein's gravity, for a spherical body, it is \(\Omega_{\rm K}^2 = \omega_z^2 > \omega_r^2\ ,\) i.e. non-circular orbits are not closed. In addition, for circular orbits with radii smaller than \(6GM/c^2\ ,\) it is \(\omega_r^2 < 0\ ,\) which indicates the dynamical instability of these orbits. We describe this and other aspects of the black hole gravity that are most relevant to the accretion disc physics in sub-section The black hole gravity of this Scholarpedia article.

Paczynski and Wiita (1980) realised that by a proper guess of an artificial Newtonian gravitational potential, \(\Phi = -GM/(r - r_G)\) (with \(r_G = 2GM/c^2\)), one may accurately describe in Newton's theory the relativistic orbital motion, and in particular the existence of ISCO. Paczynski's model for the black hole gravity became a very popular tool in the accretion disc research. It is used by numerous authors in both analytic and numerical studies. Effects of special relativity have been added to Paczynski's model by Abramowicz et al.(1996), and a generalization to a rotating black hole was done by e.g. Karas and Semerak (1999). Newtonian models for rotating black holes are cumbersome and for this reason not widely used, see Abramowicz (2009).

Viscous processes, with the timescale \(t_{\rm vis}\)

Despite the fact that the crucial role of accretion power in quasars and other astrophysical objects was uncovered already forty years ago by Salpeter and Zeldovich, several important aspects of the very nature of accretion discs are still puzzling. One of them is the origin of the viscous stresses. Balbus and Hawley recognised in 1991 that, most probably, viscosity is provided by turbulence, which originates from the magneto-rotational instability. The instability develops when the matter in the accretion disc rotates non-rigidly in a weak magnetic field. There is still no consensus on how strong the resulting viscous stresses are and how exactly they shape the flow patterns in accretion discs. A great part of our detailed theoretical knowledge on the role of this source of turbulence in accretion disc physics comes from numerical supercomputer simulations. The simulations are rather difficult, time consuming, and hardware demanding. Due to mathematical difficulties, in analytic models one does not directly implement a (small scale) magnetohydrodynamical description, but describes the turbulence (or rather the action of a small scale viscosity of an unspecified nature) by a phenomenological "alpha-viscosity prescription" introduced by Shakura and Sunyaev: the kinematic viscosity coefficient is assumed to have the form

\[\nu = \alpha H V\ ,\]

where \(\alpha =\,\,\)const is a free parameter, \(H\) is a length scale (usually the pressure scale), and \(V\) is a characteristic speed (usually the sound speed). There are several versions of this prescription, the most often used assumes that the viscous torque \(t_{r\phi} = \alpha P\) is proportional to a pressure (either the total, or the gas pressure). The rate of viscous dissipation of energy is

\[q = \nu \Sigma (r \frac{d\Omega}{dr})^2\]

Figure 9: Development of the MIR instability in a Polish doughnut, from numerical simulations by J. Hawley.
Credit: J.A. Font, Numerical Hydrodynamics and Magnetohydrodynamics in General Relativity
There is a disagreement between experts on the viscosity prescription issue: some argue that only the hydromagnetic approach is physically legitimate and the alpha prescription is physically meaningless, while others stress that at present the magnetohydrodynamical simulations have not yet sufficiently maturated to be trusted, and that the models that use the alpha prescription capture more relevant physics. All the detailed comparisons between theoretical predictions and observations performed to date were based on the alpha prescription.

Thermal processes, with the timescale \(t_{\rm the}\)

Gravitational and kinetic energy of matter falling onto the central object is converted by dissipation to heat. Heat is partially radiated out, partially converted to work on the disc expansion and (in the case of BH accretion) partially lost inside the hole. The efficiency of accretion disc \(\eta\) is defined by \(L = \eta {\dot M}c^2\ ,\) where \(L\) is the total luminosity (power) of the disc radiation. Sołtan gave a strong observational argument, confirmed and improved later by other authors, that the efficiency of accretion in quasars is \(\eta \approx 0.1\ .\) Note that the efficiency of thermonuclear reactions inside stars is about \(\eta \approx 0.007\). The theoretically predicted efficiency of geometrically thin and optically thick Shakura-Sunyaev accretion disc around a black hole is \(\eta \ge 0.1\ .\) Thus, Shakura-Sunyaev accretion discs could explain the energetics of the "central engines" of quasars, which are the most efficient steady engines known in the Universe. Other types of accretion discs models (like ADAFs and slim discs) are called the "radiatively inefficient flows" (RIFs) because they are radiatively much less efficient.

The energy budget may also include rotational energy that could be tapped from the central object. In the black hole case, this possibility was described in a seminal paper by Blandford and Znajek. The Blandford-Znajek process is an electromagnetic analogy of the well-known Penrose process. Some of its aspects are not yet rigorously described in all relevant physical and mathematical details, and some remain controversial. It is believed that the Blandford-Znajek process may power the relativistic jets.

Analytic models of accretion discs

General remarks

Analytic models describe extreme geometries, matter supplies and optical depths

Non-linear, coupled partial differential equations of radiative viscous hydrodynamics (or magnetohydrodynamics) that describe physics of accretion discs are too complex to be exactly solved analytically in the general case. Usually, analytic models assume that the accretion is stationary and axially symmetric. For such discs, useful approximate solutions exist in extreme cases corresponding to the following three fundamental divisions (as shown in Figure 1):

  • Geometry: vertically "thin" versus "thick" discs

  • Mass supply rate: "sub" versus "super" Eddington accretion rate

  • Optical depth: "opaque" versus "transparent" discs

Subsection Thin discs covers accretion disc models with H/r < 1 that includes Shakura-Sunyaev thin discs, slim discs and ADAFs.
Subsection Thick discs covers accretion disc models with H/r > 1 that includes Polish doughnuts and ion tori.

Figure 10: The main types of analytic accretion disc models in the parameter space of different geometries (i.e. vertical thickness), matter supplies (i.e. accretion rates) and optical depths. Credit: Aleksander Sadowski (2009)

Extreme geometries: vertically "thin" and "thick" accretion discs

A thin discs has its "vertical" (i.e. across the disc plane) extension much smaller than its "radial" (along the plane) extension, \(H \ll r\ .\) This means that the disc structure depends mostly on the radial coordinate \(r\) and may be described by ordinary differential equations. Thick discs have toroidal shapes with \(H \approx r\ .\) In this case, the analytic solution is possible because simplifying assumptions concerning mostly physics. Detailed models of thin and thick discs are described in the sub-sections of this Scholarpedia article: Thin discs, Thick discs.

Extreme mass supply : "sub" and "super" Eddington accretion rates

Figure 11: Radio maps of SS433, a source containing a super-Eddington accretion disc. SS433 may be a Galactic prototype of the ultraluminous X-ray sources (ULXs) found in other galaxies.

The accretion rate is defined as the instantaneous mass flux through a spherical surface \(r =\,\,\)const inside the disc. In non-stationary accretion discs accretion rate depends on both time and location, but in stationary disc models with no substantial outflows (no strong winds) it is

\[{\dot M}(r, t) = const.\]

Accretion discs may be divided into two classes, depending on whether accretion rate is much smaller than, or comparable to the characteristic Eddington accretion rate, that depends only on the mass of the central accreting object \(M\ ,\)

\[ {\dot M}_{\rm Edd} = \frac{L_{\rm Edd}}{\eta c^2} = 1.5 \times 10^{17} \, ({M/M_{\odot}}) \, {\rm g/s}. \]

\(M_{\odot} = 2 \times 10^{33}\, {\rm g}\) denotes the mass of the Sun, and \(L_{\rm Edd}\) is the Eddington luminosity (radiation power), familiar from the theory of stellar equilibria: at the surface of a star shining at the Eddington rate, the radiation pressure force balances the gravity force (see also the introduction of this Scholarpedia article).

The figure on the left shows radio maps of SS433, a well-known Galactic object with a super-Eddington accretion disc.

A rather common belief that a black hole cannot accrete at a rate higher than the Eddington limit is wrong. In particular, the Eddington rate is not a limit for the mass growth rate of a black hole due to accretion, \(dM/dt\ .\) It could be that \(dM/dt \gg {\dot M}_{\rm Edd}\ .\) This is relevant for modelling the cosmological evolution of black holes.

Extreme optical depth: "opaque" and "transparent" accretion discs

Note: a more detailed discussion of the subject presented in this Section is given in Narayan & Yi (1995). Optical depth in vertical direction \(z\) is approximated by \(\tau = \Sigma \kappa H\ .\) Here \(\kappa\) is the opacity coefficient, and \(\Sigma\) is the surface density, i.e. vertically integrated density.

Opaque discs (\(\tau \gg 1\)): Such discs are not very hot, the temperature is much less than the virial temperature, \(T \ll T_{\rm{vir}} = \frac{2}{3}\frac{G M m_p}{k_B r}\ .\) Simple (and often used) analytic models approximate the flux emitted locally (at a fixed radius \(r\)) from the disc surface by the "diffusive" black body formula, \(f = f(r) = {8\sigma T^4}/{3H\tau}\ .\) In calculating spectra, the total flux from whole surface of accretion disc is (roughly) approximated by the Planck formula,

\[ F_{\nu}=4\pi \frac{\nu^3 \cos i}{c^2 d^2}\int^{r_{out}}_{r_{in}}\frac{r}{\exp[h\nu/kT(r)]-1}dr\ ,\]

where \(d\) and \(i\) are the distance and inclination angle to the rotation axis, respectively, as seen by an observer. More advanced models solve (approximately) radiative transfer equation in the vertical direction, considering dependence on the radiation frequency \(\nu\ .\)

Transparent discs (\(\tau \ll 1\)): Such discs have relatively high-temperatures and low-densities. Bremsstrahlung, synchrotron and Compton radiative processes are most relevant,

\[f = f_{\rm{br}} + f_{\rm{br,C}} + f_{\rm{syn}} + f_{\rm{syn,C}}\ .\]

They cool down the electrons in the gas much more efficiently than the ions, and therefore a temperature separation between electrons \(T_e\) and ions \(T_i\) is expected. Radiative cooling by Bremsstrahlung,

\[f^-_{\rm{br}} = f_{ei} + f_{ee}\]

is given by,

\[f_{ei} = n_e n_i c \sigma_T \alpha_f m_e c^2 F_{ei}(\theta_e) \,\, {\rm erg \,\, cm^{-3} \,\, s^{-1}},\]

where \(F_{ei}(\theta_e<1) = 4 (\frac{2 \theta_e}{\pi^3})^{1/2} (1 + 1.781 \theta_e^{1.34})\ ;\) \(F_{ei}(\theta_e>1) = \frac{9 \theta_e}{2\pi}[\ln(1.123 \theta_e + 0.48) + 1.5]\)


\[f_{ee} = n_e^2 c r_e^2 \alpha_f m_e c^2 F_{ee}(\theta_e) \,\, {\rm erg \,\, cm^{-3} \,\, s^{-1}},\]

where \(F_{ee}(\theta_e<1) = \frac{20}{9 \pi^{1/2}}(44-3\pi^2)\theta_e^{3/2} (1+1.1\theta_e^2 - 1.25\theta_e^{5/2})\ ;\) \(F_{ee}(\theta_e>1) = 24 \theta_e [\ln(0.5616 \theta_e) + 1.28]\ .\)

A simple approximation inverse Compton scattered Bremsstrahlung emission writes

\[f_{\rm{br,C}} = 3\eta_1\{\frac{1}{3}(1-\frac{x_c}{\theta_e})-\frac{1}{\eta_3+1}[(\frac{1}{3})^{\eta_3+1}-(\frac{1}{3\theta_e})^{\eta_3+1}]\} f_{\rm{br}}, \]

where \(n_e\) and \(n_i\) are the number densities of electrons and ions, \(\sigma_T\) is the Thomson cross-section and \(\alpha_f\) the fine structure constant, \(m_e\) and \(r_e = e^2/m_ec^2\) are the electron's mass and radius and \(c\) the speed of light. \(F_{ee}(\theta_e)\) and \(F_{ei}(\theta_e)\) are the radiation rate functions, given by the dimensionless electron temperature \(\theta_e = kT_e/m_ec^2\, \). The Compton enhancement factor \(\eta = \eta_1 + \eta_2 + ({x}/{\theta_e})^{\eta_3}\) is given by \(\eta_1 = P(A-1)/(1-PA)\), \(\eta_1 = -3^{-\eta_3} \eta_1\) and \(\eta_3 = - 1 - \ln P / \ln A\ ,\) where \(x = h\nu/m_ec^2\, \) factor \(P\) is the probability that a photon scatters and \(A\) is the mean energy amplification factor by that photon and \(x_c = h\nu_c/m_e c^2\ .\) If a magnetic field \(B\) is present, there is also radiative cooling by synchrotron emission

\[f_{\rm{syn}} = \frac{2\pi}{3c^2}kT_e(r)\frac{d\nu_c^3(r)}{dr}\,, ~~ \nu_c = \frac{3 e B}{4 \pi m_e c} \theta_e^2 x_M,\]

\[f_{\rm{syn,C}} = [\eta_1 - \eta_2(\frac{x_c}{\theta_e})^{\eta_3}] f_{\rm{syn}},\]

where the coefficient \(x_M\) must be numerically calculated from a relativistic Maxwellian distribution of electrons. For a more thorough study of Comptonisation see, e.g., Coppi & Blandford (1990).

Thin discs

Expansion in the "smallness" parameter. The "global" parameters of a thin disc: mass, accretion rate, viscosity

Most of the analytic and semi-analytic accretion disc models assume that the disc is stationary and axially symmetric. "Thin" disc models assume in addition that the vertical extension of the disc is small in a general sense that in cylindrical coordinates \(r, z, \phi\) the disc surface is given by \(z_{\pm} = H_{\pm}(r)\) and \(\max(|H/r|) \equiv \epsilon \ll 1\ .\) Here \(z = 0\) describes the location of the accretion disc plane. In the spherical coordinates the plane is given by \(\cos \theta = 0\ ,\) and the condition of the small vertical extension by \(\cos \theta \ll 1\) everywhere inside the disc.

The thin disc models are based on expanding the hydrodynamic (or MHD) equations in powers of \(\epsilon\ .\) The expansion procedure is not unique, and depends on some extra physical assumptions made. It leads to equations of the general form \(A_0^i\epsilon^0 + A_1^i\epsilon^1 + A_2^i\epsilon^2 + ... = 0\ .\) For most models, the resulting set of equations, \(A^i_k = 0\ ,\) consists of a number of coupled, linear first-order ordinary differential equations (containing \(d/dr\) derivatives) and a few non-linear algebraic equations. Usually, the integration constants may be associated with (and calculated from) the three "global" parameters of thin disc models that are the mass of the central accreting object \(M\ ,\) the accretion rate \({\dot M}\ ,\) and the viscosity parameter \(\alpha\ .\)

The "standard thin", "ADAF" and "slim" disc equations

Newtonian hydrodynamic models of stationary and axially symmetric, thin accretion discs are described by equations similar to (or equivalent to) the 12 equations given in the table below. A "model" should give each of the 12 unknown quantities, for example the matter density \(\rho\ ,\) as a function of the radius \(r\ ,\) and the three model parameters \(\rho = \rho(r, M, {\dot M}, \alpha)\ .\)

      Unknown functions
(01) \(\rho\,V\frac{dV}{dr} = \rho (\Omega^2 - \Omega_K^2 )\,r - \frac{dP}{dr}\) radial balance of forces

\(\rho\) density of matter
(02) \({\dot M}\left[ \frac{dU}{dr} + P\,\frac{d}{dr}\left(\frac{1}{\rho}\right)\right] = 4\pi r^2 H(\tau_{r \phi}) \frac{d\Omega}{dr} + 4\pi r\,F\) energy conservation

\(V\) radial velocity
(03) \({\dot M} = 4\pi\,r\,H\rho\,V\) mass conservation

\(\Omega\) angular velocity of matter
(04) \({\dot M}\left(j - j_{0}\right) = 4\pi\,r^2\,H(\tau_{r\phi}) - X_{0}\) angular momentum conservation

\(P\) pressure
(05) \(F = \frac{acT^{4}}{[\kappa\,\rho\,H]}\) vertical radiative transfer

\(U\) internal energy (thermodynamics)
(06) \(\frac{c_s}{\Omega_K\,r} = \frac{H}{r}\) vertical momentum balance

\(H\) vertical thickness
(07) \((\tau_{r\phi}) = \rho[\alpha\,c_s\,H]r \frac{d\Omega}{dr}\) "viscous" torque \(\alpha\)prescription

\((\tau_{R\phi})\) viscous torque
(08) \(P = P(\rho, T)\) equation of state

\(F\) flux of radiation
(09) \(U = U(\rho, T)\) equation of state (internal energy)

\(T\) temperature
(10) \(c_s = c_s(\rho, T)\) equation of state (sound speed)

\(j\) angular momentum
(11) \(\kappa = \kappa (\rho, T)\) opacity

\(\kappa\) opacity coefficient
(12) \(j = \Omega\,r^2\) angular momentum and angular velocity \(c_s\) sound speed
NOTE: these equations are valid for the standard Shakura-Sunyaev discs, for ADAFs and for slim discs. In the case of the standard Shakura-Sunyaev discs further assumptions are made, which transform all the equations to the algebraic ones. Specifically, in equation (01) one puts \(dV/dr = 0 = dP/dr\) which leads to \(\Omega^2 = \Omega^2_{\rm{K}} = GM/r^3\ .\) In equation (02) one puts \(dU/dr = 0 = d\rho/dr\) which leads, after some manipulations involving other equations, to

\[F = \frac{3GM\dot{M}}{8\pi r^3} \left(1 - \sqrt {\frac{r_0}{r}}\right)\]

This is the famous Shakura-Sunyaev flux formula. Note also that the gravitation field of the central object enters the above Newtonian equations only through the Keplerian angular velocity \(\Omega_{\rm{K}}(r)\ .\) In the general relativistic version of (01)-(12) the gravity (i.e. the spacetime curvature) enters also through components of the metric tensor \(g_{\mu\nu} = g_{\mu\nu}(r)\) the equatorial plane. The Kerr geometry version of (01)-(12) was written first by Lasota (1994), and later elaborated by Abramowicz, Chen, Granath and Lasota (1996); see also Sadowski (2009). However, equations (01)-(12) in their form above are often used to model the black hole accretion discs. This is possible because of a brilliant discovery by Paczynski of the Newtonian model for the black hole gravity.

Particular models make several additional simplifying assumptions. For example, several models assume that \(\Omega(r) = \Omega_{\rm{K}}(r)\ ,\) with \(\Omega_{\rm{K}}(r)\) being the Keplerian angular velocity, which is known since the gravitational field of the central object is known (for a spherical body with the mass \(M\) Newton's theory yields \(\Omega_{\rm{K}}(r) = (GM/r^3)^{1/2}\)). Note, that in this case the derivative \(d\Omega/dr\) that appears in equations (02) and (07) becomes a known function of \(r\ .\) Equation (07) postulates the form of the "viscous" stress \((\tau_{r\phi})\) in therms of an ad hoc ansatz that introduces the dimensionless \(\alpha\)-viscosity. Note that the quantity that appears in square bracket is called in hydrodynamics the "kinematic viscosity". The original Shakura-Sunyaev ansatz postulated \((\tau_{r\phi}) = \alpha\,P\ .\) Equation (05) gives the flux of radiation in (a very rough) diffusion approximation. Note that the quantity in square brackets in this equation is the optical depth, \(\tau = [\kappa\,\rho\,H]\) in the vertical direction. The equation is valid only if \(\tau \gg 1\ ,\) and if \(\tau < 1\) non-thermal radiative processes should be considered, and equation (05) replaced by \(F = F(\rho, T)\ .\)

In equation (04), \(j_{0}, X_{0}\) are the angular momentum and the viscous torque at some undefined radius \(r_{0}\ .\) In the black hole accretion discs models, it is customary to take \(r_{0} = r_G\) (= black hole horizon radius) because the viscus torque at the horizon vanishes. Then, \(j_{0}\) is the (a priori unknown) angular momentum of matter at the horizon. With respect to first order derivatives, equations (01)-(12) form a linear system that may be solved for each derivative. For \(dV/dr\) this gives, \[ \frac{dV}{dr} = \frac{N(r, \rho, V, \Omega, ...)}{V^2 - C^2_s}.\] Any black hole accretion flow must be transonic, i.e. somewhere it must pass the sonic radius \(r_s\ ,\) where \(V(r_s) = c_s(r_s)\ .\) In order that \(dV/dr\) and all other derivatives are non-singular there, it must be, \[N(r_s, \rho, V, \Omega, ...) = 0.\] The above sonic point regularity condition makes the system (01)-(12) over constrained, i.e. an eigenvalue problem, with the eigenvalue being the angular momentum at the horizon, \(j_{0}\ .\)

Analytic models describe black hole accretion discs down to a certain "inner edge" \(r_{in}\) which locates close to the central accreting object. The inner edge is a theoretical concept introduced for convenience, because at \(r \approx r_{in} \approx r_s\) the accretion flow changes its character. In the case of the black hole accretion, the change goes from almost circular orbits to almost radial free fall. It is therefore convenient to separately model the two regions\[r > r_{in}\] where matter moves on circular orbits, and \(r < r_{in}\) where matter free falls. Of course, in reality the situation is more complicated, as the change of the flow character occurs smoothly in an extended region on both sides of \(r_{in}\ .\) For black hole accretion the inner disc edge lies between the marginally bound radius and the innermost stable circular orbit, \(r_{\rm{MB}} \le r_{in} \le r_{\rm{ISCO}}\,\). For very efficient Shakura-Sunyaev discs, \(r_{in} \approx r_{\rm{ISCO}}\ ,\) while for RIFs \(r_{in} \approx r_{\rm{MB}}\ .\) For stellar accretion, \(r_{in}\) is located near the surface of the star and the flow there is described by a boundary layer model. For more details on the inner edge of a thin disc, see Abramowicz et al. (2010) and references quoted there.

The "standard" model: Shakura-Sunyaev

Characteristic, applications, references

Analytic formulae (solution)

Axially symmetric, stationary, local analytic model. Explicit formulae give all physical characteristics in terms of \(M, {\dot M}, \alpha\) and \(r\ .\) It is geometrically thin in the vertical direction \(H/r < 1\ ,\) and has a disc-like shape. Accretion rate is very sub-Eddington. Opacity is very high. The gas goes down on tight spirals, approximated by circular, free (Keplerian, geodesic) orbits. For black hole and (very compact) neutron star the inner edge locates at ISCO. High luminosity, high efficiency of radiative cooling. Electromagnetic spectra is not much different from that of a sum of black bodies. Alpha viscosity prescription assumed. Diffusion approximation for radiative transfer used. It is dynamically stable. When the gas is cold and radiation pressure negligible, it is also thermally and viscously stable, otherwise it is unstable in both respects.

Applications: YSOs, CVs, LMXRB, AGNs. The best known and studied theoretical model.

Standard reference: Shakura, Sunyaev (1974) , one of the most often quoted papers in modern astrophysics (quotation counts).

Similar ideas: Pringle, Rees (1972); Lynden-Bell, Pringle (1974)

Fully relativistic version (Novikov, Thorne 1974);  for the detailed description see: Page, Thorne (1974).  Recent application to spectral fits: Shafee et al. (2006); Middleton et al. (2006).

Recommended review: Pringle (1981).

Different analytic solutions are known for cases in which the total pressure is either dominated by gas or radiation pressure, and for the opacity described either by Kramers' law, or electron scattering. However, the most important formula that gives the locally emitted flux of radiation does not depend on these physical conditions, and is universally given (in Newton's theory) by,
\( F = \frac{3GM\dot{M}}{8\pi r^3} \left(1 - \sqrt {\frac{r_0}{r}}\right)\)
The Newtonian formulae below correspond to the case of gas pressure and Kramers' law opacity law

\(H = 1.7\times 10^8\alpha^{-1/10}\dot{M}^{3/20}_{16} m_1^{-3/8} r^{9/8}_{10}f^{3/5} {\rm cm}\)

\(T = 1.4\times 10^4 \alpha^{-1/5}\dot{M}^{3/10}_{16} m_1^{1/4} r^{-3/4}_{10}f^{6/5}{\rm K}\)

\(\rho = 3.1\times 10^{-8}\alpha^{-7/10}\dot{M}^{11/20}_{16} m_1^{5/8} r^{-15/8}_{10}f^{11/5}\rm{g/cm^3}\)

Here \(T\) and \(\rho\) are the mid-plane temperature and density respectively. \(\dot{M}_{16}\) is the accretion rate, in units of \(10^{16} \rm{g/s}\ ,\) \(m_1\) is the mass of the central accreting object in units of a solar mass, \(r_{10}\) is the radial location in the disc, in units of \(10^{10}{\rm{cm}}\ ,\) and \(f = [1-(r_0/r)^{1/2}]^{1/4}\ ,\) where \(r_0\) is the inner radius of the disc.

Notes: The flux formula in the box is (probably) the most often used one in the accretion disc research. It shows that the flux does not depend on \(\alpha\ ,\) the viscosity parameter. Other physical quantities depend on \(\alpha\) rather weakly. This which is a very fortunate feature of the Shakura-Sunyaev model, as the \(\alpha\)-viscosity prescription is assumed ad hoc and not derived from the first principles.

Specific versions (and modifications) of the Shakura-Sunyaev models

S-curves and the thermal-viscous instability: The standard accretion discs are known to be subject to thermal-viscous instability due to the partial hydrogen ionisation. As a result of this instability, the disc cycles between two states: a hot and mostly ionised state with a large local accretion rate and a cold, neutral state with a low accretion rate. This instability was originally proposed to explain the large-amplitude luminosity variations observed in cataclysmic variables (Smak 1982; Meyer & Meyer-Hofmeister 1982). It is also believed that the same mechanism is responsible for the eruptions in soft X-ray transients (see, e.g. Cannizzo et al. 1982; Dubus et al. 2001; and Lasota 2001 for a review). The H ionisation instability was also shown to operate in some discs around supermassive black holes in active galactic nuclei (Lin & Shields 1986; Clarke 1988; Mineshige & Shields 1990; Siemiginowska et al. 1996), but not necessarily on a global scale (see Menou & Quataert 2001; Janiuk et al. 2004). The characteristic timescales of cycle activity scale roughly with the mass of a compact object (Hatziminaoglou et al. 2001). Therefore, the observed cycle timescale of the order of years in binaries translates into thousands to millions of years in galaxies that harbour a supermassive black hole.

Warped and precessing discs: A geometrically thin, optically thick accretion disc is unstable to self-induced warping when illuminated by a sufficiently strong central radiation source (Pringle 1996; Maloney et al. 1996; see also Petterson 1977). The instability is important for the standard (not advectively dominated) discs around neutron stars and black holes in the X-ray binaries, in active galactic nuclei (Pringle 1997), and in particular in the "maser" galaxy NGC 4258 (Maloney et al. 1996). For discs around less compact objects, where efficiency is orders of magnitude smaller, steady discs are predicted to be stable. The warped discs are important in some models of the (~kilohertz) coherent oscillations (QPOs) observed in neutron star and black hole X-ray binaries. In the Schwarzschild metric, the damping or excitation of g-modes with azimuthal dependence \(\exp(i\,m\phi)\) has been considered by Kato (2003, 2004, 2005, 2007), who found that m = 0, 1 modes undergo resonant amplification when interacting with a nonrotating one-armed (m = 1) stationary warp, assumed to be present in the accretion disc. Description of the warp instability here is based on Armitage & Pringle 1997. See a critical examination of the idea e.g. in Ivanov & Papaloizou 2008.


Characteristic, applications, references

Radial distributions

Sub-Eddington accretion, very small opacity. ADAFs are cooled by advection (heat captured by moving matter) rather than by radiation. They are very radiatively inefficient, geometrically extended, similar in shape to a sphere (or a "corona") rather than a disc, and very hot (close to the virial temperature). Because of their low efficiency, ADAFs are much less luminous than the Shakura-Sunyaev thin discs. ADAFs emit a power-law, non-thermal radiation, often with a strong Compton component. For black hole and (very compact) neutron star the inner edge at a radius smaller than ISCO. Dynamically, thermally and viscously stable.

Applications: mostly LMXRB, AGNs, with good fits to observed spectra.

Numerical 1.5D (vertically integrated) stationary transonic models (in Kerr): Abramowicz et al. (1996); Narayan et al. (1997); Popham, Gammie (1998)  Numerical 2D non-stationary models (in Paczynski-Wiita): Igumenshev et al. (1996).

Most influential paper, describing a Newtonian, self-similar, stationary, axially symmetric analytic, model: Narayan, Yi (1994). Immediate follow-up by the same group: Narayan, Yi (1995); Narayan, Mahadevan (1995); Narayan et al. (1996).

The idea mentioned first time: Ichimaru (1987), see also: Rees et al. (1982); Abramowicz et al. (1995).

Recommended review: Narayan, McClintock (2008).


The profiles of temperature, optical depth, ratio of scale height to radius, and the advection factor of a hot one-temperature accretion solution (solid lines). The parameters are\[M = 10\,M_{\odot}\ ,\] \({\dot M} = 10^{-5}{\dot M}_{\rm{Edd}}\ ,\) \(\alpha = 0.3\ ,\) \(\beta = 0.9\ .\) The outer boundary conditions are \(R_{out} = 10^3R_s\ ,\) \(T = 10^9\)[K], \(v/c_s = 0.5\ .\) For simplicity, effects associated with outflow or convection are not taken into account. The two-temperature solutions with the same parameters and \(\delta = 0.5\) (dashed lines) and 0.01 (dot-dashed lines) are also shown for comparison. Figure credit: Yuan, Taam, Xue, Cui (2006)

Specific versions (and modifications) of the ADAF models

Convection: CDAFs: The Convection Dominated Accretion Flow (CDAF) models an ADAF with convective eddies in which gas can be trapped. This leads to radial modifications of the accretion rate.

Outflows: ADIOS The Adiabatic Inflow-Outflow Solution (ADIOS) accounts for the loss of gas due to outflow. In this model the accretion rate decreases with decreasing radius.

Magnetized ADAFs

Slim discs

Characteristic, applications, references

The local flux

Nearly Eddington accretion. Large opacity. Cooled by radiation and advection, i.e. heat trapped in matter and transported towards the central accretor. Radiatively much less efficient than the standard Shakura-Sunyaev discs. \(H/R\) only slightly less than one. For black hole and (very compact) neutron star the inner edge at a radius smaller than ISCO. Dynamically, thermally and viscously stable. Rotation differs (slightly but importantly) from the Keplerian one. The pressure gradient along the disc plane direction is dynamically important. Slim discs models are described by a set of ordinary differential equations, and one must explicitly solve the eigenvalue problem connected with the regularity condition at the sonic radius (which does not coincide with the ISCO).

Applications: mostly LMXRB, AGNs, ULX. For detailed spectral fits see e.g.:

  • Most sophisticated slim disc models (in the Kerr geometry): Sadowski (2009)

Figure on the right shows the local flux of radiation for different mass accretion rates and the black hole spins. Each subplot contains six solid lines for the following mass accretion rates (in the Eddington units): 0.01 (the thickest line), 0.1, 0.3, 0.6 and 0.9 (the thinnest line). The upper panel is for a non rotating blach hole (a = 0), the middle one for (a = 0.6) while the bottom one for a highly spinning black hole (a = 0.98). The black hole mass is \(9.4\,M_{\odot}\ .\) NOTE that a considerable amount of energy is radiated inside the ISCO; this is an effect of advection. Figure credit: Sadowski (2009)



Fully two dimensional analytic solution (stationary, axially symmetric) obtained through a mathematically exact expansion in the small parameter H/r of the equations of viscous hydrodynamics. Significant backflows in the midplane of the disc have been found.

Kluzniak, Kita (2000)numerical follow up: Umurhan et al. (2006).

Disc solutions

Branches I-III are thin disc solutions, branch IV is a thick disc solution. Lines correspond to fixed M, r, and \(\alpha\ .\) An example of each of the four branches is shown in a corresponding color: pink, blue, green, and yellow. The congruence of all branches has a critical point, corresponding to \( \alpha = \alpha_{crit}\ . \) In different places of the parameter space, the cooling is dominated by black body radiation, bremsstrahlung, Compton losses, pair production, or by advection, as indicated by arrows. Figure adapted from Björnsson et al. (1996).

Branch I (blue): Shakura-Sunyaev (gas pressure) + Shakura-Sunyaev (radiation pressure) + Slim.

Branch II (green): Shakura-Sunyaev (gas pressure) + SLE

Branch III (yellow): SLE + ADAF.

Branch IV (pink): Polish doughnut (thick disc, see next section).

The four branches of analytic models of accretion discs.


The different types of thin discs may coexist radially. Microquasaes display distinct spectral states. In order of increasing luminosity these are the quiescent state, low state, intermediate state, high state, and very high state. Narayan with collaborators presented a model of accretion flows around black holes that unifies most of these states. At low mass accretion rates, the inner ADAF zone in the model radiates extremely inefficiently, and the outer thin disc is restricted to large radii. The luminosity therefore is low, and this configuration is identified with the quiescent state. For larger accretion rates the radiative efficiency of the ADAF increases rapidly and the system becomes fairly luminous. The spectrum is very hard and peaks around 100 keV. This is the low state. For still greater rates, the ADAF progressively shrinks in size, the transition radius decreases, and the X-ray spectrum changes continuously from hard to soft. This is the intermediate state. Finally, the inner ADAF zone disappears altogether and the thin accretion disc extends down to the marginally stable orbit. The spectrum is dominated by an ultrasoft component with a weak hard tail. This is the high state.

Thick discs

Thick discs: assumptions

For "thick discs" models of accretion discs one assumes that:

  • Matter distribution is stationary and axially symmetric, i.e. matter quantities such as density \(\epsilon\) or pressure \(P\) are independent on time \(t\) and the azimuthal angle \(\phi\ .\)
  • Matter moves on circular trajectories, i.e. the four velocity has the form \(u^i = [u^t, u^{\phi}, 0, 0]\ .\) The angular velocity is defined as \(\Omega = u^{\phi}/u^t\ ,\) and the angular momentum as \(\ell = - u_{\phi}/u_t\ ,\)
  • \(t_{\rm dyn} \ll t_{\rm the} < t_{\rm vis}\ ,\) with \(t_{\rm dyn}\) being the dynamical timescale in which pressure force adjusts to the balance of gravitational and centrifugal forces, \(t_{\rm the}\) being the thermal timescale in which the entropy redistribution occurs due to dissipative heating and cooling processes, and \(t_{\rm vis}\) being the viscous timescale in which angular momentum distribution changes due to torque caused by dissipative stresses. Mathematically, this is equivalent to assume the stress-energy tensor in the form, \(T^i_{~\nu} = u^{\mu}\,u_{\nu}\,(P + \epsilon) - \delta^{\mu}_{~\nu}\,P\ .\)

Using this form of the stress-energy tensor, Abramowicz et al. 1978 have derived from the equilibrium condition \(\nabla_{\mu}\,T^{\mu}_{~\nu} = 0\) the relativistic "Euler" equation,

\[ \frac{\nabla_{\mu} P}{\epsilon + P} = \frac{\nabla_{\mu}\,g_{tt} + 2\Omega\,\nabla_{\mu}\,g_{t\phi} + \Omega^2\,\nabla_{\mu}\,g_{\phi\phi}}{g_{tt} + 2\Omega\,g_{t\phi} + \Omega^2\,g_{\phi\phi}} = \nabla_{\mu}\ln A + \frac{\ell\,\nabla_{\mu}\Omega}{1 - \ell\,\Omega}, ~~~{\rm with}~~~A^2(r, \theta) = \frac{1}{g_{tt}(r,\theta) + 2\Omega\,g_{t\phi}(r,\theta) + \Omega^2\,g_{\phi\phi}(r,\theta)}. \]


Equipressure surfaces: analytic solution in a general case

The "equipressure" surfaces are defined by an implicit condition \(P(r, \theta) = const\ ,\) which may be solved to get the explicit form \(\theta = \theta(r)\ .\) Then, (3.2:1) implies that the function \(\theta(r)\) obeys,

\[ \frac{d\theta}{dr} = - \frac{\partial_{r}\,g_{tt} + 2\Omega\,\partial_{r}\,g_{t\phi} + \Omega^2\,\partial_{r}\,g_{\phi\phi}} {\partial_{\theta}\,g_{tt} + 2\Omega\,\partial_{\theta}\,g_{t\phi} + \Omega^2\,\partial_{\theta}\,g_{\phi\phi}} = - \frac{\partial_{r}\,g^{tt} - 2\ell\,\partial_{r}\,g^{t\phi} + \ell^2\,\partial_{r}\,g^{\phi\phi}} {\partial_{\theta}\,g^{tt} - 2\ell\,\partial_{\theta}\,g^{t\phi} + \ell^2\,\partial_{\theta}\,g^{\phi\phi}}. \]


If \(\ell = \ell(r, \theta)\) or, which is equivalent, \(\Omega = \Omega(r, \theta)\) are known functions, the equation (3.2:2) for the equipressure surfaces takes the form of a standard ordinary differential equation \(d\theta/dr = f(r, \theta)\) with known rhs, and it may be directly integrated. This has been done first by Jaroszynski et al. (1980) and then by several other authors. In the Figure below, we show an example from a recent paper by Lei et al. (2008) who assumed the angular momentum distribution in the form that depends on three constant parameters (\(\eta, \beta, \gamma\)),

\[ {\ell}(r, \theta) = {\ell}_0\left\{ \frac{{\ell}_K(r)}{{\ell}_0}\right\}^{\beta}(\sin\theta)^{2\gamma} ~\left({\rm for}~~ r \geq r_{\rm ISCO}\right) ,~~ {\ell}(r, \theta) = {\ell}_{\rm ISCO}(\sin\theta)^{2\gamma} ~\left({\rm for}~~ r < r_{\rm ISCO}\right). ~~{\rm Here}~~ {\ell}_0 \equiv \eta\,{\ell}_K(r_{\rm ISCO}) ~~{\rm and}~~ {\ell}_{\rm ISCO} = {\ell}_0\left\{\frac{{\ell}_K(r_{\rm ISCO})}{{\ell}_0}\right\}^{\beta}. \]


Fig. 1 Equipressure surfaces for a non rotating black hole, and with angular momentum distribution (3.2:3) with \(\gamma = 0.00\ ,\) \(\eta = 1.085\ ,\) \(\beta = 0.00\ .\) Angular momentum is constant everywhere in space.


Fig. 2 Equipressure surfaces for a non rotating black hole, and with angular momentum distribution (3.2:3) with \(\gamma = 0.50\ ,\) \(\eta = 1.085\ ,\) \(\beta = 0.00\ .\) Angular momentum is constant radially, but it changes along the polar angle.


Fig. 3 Equipressure surfaces for a non rotating black hole, and with angular momentum distribution (3.2.3) with \(\gamma = 0.18\ ,\) \(\eta = 1.085\ ,\) \(\beta = 0.9\ ,\) compared with global MHD numerical simulations by Fragile et al. (2007), time averaged over a period corresponding to the orbital period at \(25r_G\) (color-coded).


Fig. 4 Equipressure surfaces for a non rotating black hole, and with angular momentum distribution (3.2.3) with \(\gamma = 0.00\ ,\) \(\eta = 1.085\ ,\) \(\beta = 0.99\ .\) Angular momentum does not depend on the polar angle, its radial distribution is almost Keplerian.

Barytropic thick discs and the von Zeipel theorem

From equation (3.1:1) one proves (see e.g. Abramowicz et al. 1978) that for barytropic fluids \(\epsilon = \epsilon(P)\ ,\) the surfaces of constant angular velocity and of constant angular momentum coincide, i.e. \(\ell = \ell(\Omega)\ .\) This is often called "the von Zeipel theorem". The equation of state \(\epsilon = \epsilon(P)\) and the rotation law \(\ell = \ell(\Omega)\) are independent and may be separately assumed. When they are known, the analytic solution is given by,

\[ W(P) \equiv \int \frac{dP}{\epsilon(P) + P} = \ln A + \int \frac{d\Omega}{1 - \Omega\,\ell(\Omega)} \equiv \ln A(r, \theta) + F(\Omega), ~~~{\rm and}~~~\ell = \ell(\Omega) = - \frac{g_{t\phi}(r, \theta) + \Omega\,g_{\phi\phi}(r, \theta)}{g_{tt}(r, \theta) + \Omega\,g_{t\phi}(r, \theta)}. \]

The functions \(W = W(P)\ ,\) \(F = F(\Omega)\) and \(\Omega = \Omega(r, \theta)\) are explicitly known, and therefore one knows explicitly location of the equipressure surfaces \(P = P(r, \theta)\ .\) In the special (but important) case \(\ell = \ell_0 = const\ ,\) the function \(\theta = \theta(r)\) that gives the location of equipressure surfaces \(P(r, \theta) = const\) is given explicitly (for a non-rotating black hole)

\[ \sin^2\theta = \frac{\ell_0}{C_0\,r^2 + 2\,G\,M\,\left(1 - r_G/r\right)^{-1}}. ~~~{\rm Here}~~C_0=const ~~{\rm numerates~the~equipressure~surfaces}. \]

Putting \(r_G = 0\) in (3.2:5), one recovers the Newtonian formula for \(\ell = \ell_0 = const\) discs in the \(-G\,M/r\) potential.

The Roche lobe overflow

Fig. 5 Taken from Abramowicz et al. (1980) (a) At the location \(r = r_{in}\ ,\) called the "cusp", angular momentum in the disc equals the Keplerian one, \(\ell_{\rm disc}(r_{in}) = \ell_K(r_{in})\ .\) For \(r > r_{in}\) it is \(\ell_{\rm disc}(r_{in}) > \ell_K(r_{in})\) and \(d\ell/dr > 0\ ,\) and for \(r < r_{in}\) it is \(\ell_{\rm disc}(r_{in}) < \ell_K(r_{in})\) and \(d\ell/dr \approx 0\ .\) (b) The particular equipotential surface \(W = W_{in}\ ,\) called the "Roche lobe", crosses itself at the cusp. For \(r \gg r_{in}\) the surface of the disc (\(P = 0\)) coincides with the equipotential \(W = W_{S}\ .\) (c) The non-zero potential difference \(\Delta W = W_S - W_{in}\) implies that no equilibrium is possible at radii around and smaller than the cusp. Instead, there will be dynamical mass loss from the disc with the accretion rate that for a polytropic fluid \(P = K\rho^{1+1/n}\) equals (Kozlowski et al. 1978, Abramowicz 1985)

\[ {\dot M}_{in} = (2\pi)^{3/2}\frac{\Gamma(n + 3/2)}{(1 + 1/n)^n (1 + 1/2)^{n + 3/2}\Gamma(n + 3)}K^{-n}\frac{r_{in}}{\Omega_K(r_{in})}\Delta W^{n + 1}, ~~{\rm with}~~ \Gamma(m) ~~{\rm being~the~Euler~gamma~function}. \]

The mass loss (3.2:5) induced by the Roche lobe overflow self-regulates the accretion rate in the innermost part of all types of accretion discs (thin, slim, ADAF, thick) around black holes and sufficiently compact neutron stars. This self-regulated overflow has several important consequences:
  • It locally stabilises accretion discs against thermal and viscous instabilities (Abramowicz, 1981) and globally against the Papaloizou and Pringle instability (Blaes, 1987)
  • The amount of overflow, and therefore \({\dot M}_{in}\ ,\) may be modulated by global discs oscillations. In the case of neutron stars, this leads to a modulation of the luminosity of the boundary layer at the neutron star surface. Although oscillations originate in the disc, the are observed in radiation that comes from the boundary layer. This is relevant for the observed neutron star quasi periodic oscillation (QPO) (Horak et al., 2007).
  • The "runaway" instability occurs when the mass exchange between accretion disc and black hole causes the cusp to move deeper into the disc, increasing the mass lost rate \({\dot M}_{in}\ .\) This effect was suggested by Abramowicz et al. (1983), and studied by several other authors (e.g. Daigne and Font, 2004 or Montero et al., 2008). It may determine the life-time of a massive torus around a black hole, which is relevant to some models of gamma ray bursts.
Fig. 6 Taken from Igumenshchev and Beloborodov (1997). The analytic formula (3.2:5) is very accurate, as a comparison with the numerical simulations shows. (a) The numerically calculated equipressure structure close to the central black hole is remarkably similar to that calculated analytically. In particular, there is obviously a cusp there. (b) In a qualitative agreement with the analytic model, the radial fluid velocity (described by arrows) is small far away from the cusp \(r \gg r_{in}\ ,\) but large (no equilibrium) at \(r \approx r_{in}\ .\) (c) Even more impressive is the excellent quantitative agreement of the predictions of the analytic formula (3.2:5), represented by lines, with results of numerical simulations, represented by points (circles, squares and triangles).

Super-Eddington luminosity of radiation pressure supported thick discs ("Polish doughnuts")

The maximal energy available from an object with the mass \(M\) (and gravitational radius \(R_G = GM/c^2\)) is \(E_{\rm max} = M\,c^2\ .\) The minimal time in which this energy may be liberated is \(t_{\rm min} = R_G/c\ .\) Thus, the maximal power \(L_{\rm max} = E_{\rm max}/t_{\rm min} = c^5/G \equiv L_{\rm Planck} = 10^{58}\,[{\rm erg/sec}] = 10^{52}\,[{\rm Watts}]\ .\) An object with mass \(M\) has the "gravitational cross-section" \(\Sigma_{\rm grav} = 4\,\pi\,R^2_G\ .\) If radiation interacts with matter by electron scattering (with the Thomson cross section \(\sigma_T = 8\pi\,e^4/3\,m^2_e\)), its "radiation cross section" is \(\Sigma_{\rm rad} = (M/m_P)\sigma_T\ .\) M. Sikora noticed (unpublished) that the upper limit for the radiative power of an object in which gravity and radiation pressure are in equilibrium is given by the Planck power and the object gravitational and radiative cross sections,

\[ L_{\rm Edd} = L_{\rm Planck}\frac{\Sigma_{\rm grav}}{\Sigma_{\rm rad}} = \frac{4\pi\,G\,M\,m_P\,c}{\sigma_T} = 1.4 \times 10^{38}\left( \frac{M}{M_{\odot}}\right) [{\rm erg}/{\rm sec}]. \]

A "Polish doughnut" is a radiation pressure supported thick accretion discs around a central black hole. Polish doughnuts have toroidal shapes, resembling a large sphere (\(r \gg r_G\)) with a pair of empty narrow funnels along the rotation axis. The total luminosity of a Polish doughnut may be approximated as \(L/L_{\rm Edd} \equiv \lambda \approx \log (r/r_G)\ .\) The logarithm here is of a crucial importance. It prevents astrophysically realistic doughnuts (i.e. with \(r < 10^6\,r_G\ ,\)) to have highly super-Eddington luminosities. The theory predicts for such "realistic" fat tori slightly super-Eddington total (isotropic) luminosities \(\lambda \le 7\ .\) However, because the funnels have solid angles \(\Theta^2 \sim r_G/r\ ,\) radiation in the funnels may be, in principle, collimated to highly super-Eddington values \(\lambda \sim r/r_G \gg 1\ .\) This simple estimate agrees with a more detailed modelling of the Polish doughnuts radiation field by Sikora (1981) and Madau (1988) who obtained \(\lambda \ge 10^2\) for discs with \(r/r_G \sim 10^2\ .\) A typical value that follows from observational estimates for the non-blazar active galactic nuclei, e.g. by Czerny & Elvis (1987) is \(\lambda \sim 10\ ,\) but of course for blazars and other similar sources (e.g. for ULXs, if they are powered by stellar mass black holes, as argued e.g. by King (2008), it must be \(\lambda \gg 10\ .\) Fig. 7 (Credit: Madau (1988)) shows radiation from a Polish doughnut seen at different inclination angles, inclination \(0^{\circ}\) corresponds to line of sight along the funnel axis.

\[ {\dot M}_{\rm Edd} = \frac{L_{\rm Edd}}{c^2} = 1.5 \times 10^{17}\left( \frac{M}{M_{\odot}}\right) [{\rm g}/{\rm sec}], ~~~{\dot m} = \frac{\dot M}{{\dot M}_{\rm Edd}}. \]

Theoretical predictions about super-Eddington accretion rates:
  • Radiation pressure supported black hole thick accretion discs ("Polish doughnuts") have typically super-Eddington luminosities \(\lambda > 1\ .\)
  • These discs have very small accretion efficiency and therefore must have highly super-Eddington accretion rates \({\dot m} \gg 1\ .\)
  • Super-Eddington accretion does not necessarily imply strong outflows but is often accompanied by them.
The assumption that \({\dot m} = 1\) is the upper limit for the growth rate of the seed black holes, often adopted in the context of the cosmic structure formation, is false.

Stability, oscillations, variability

One of the key features observed in accretion discs is the strong and often chaotic time variability in the X-ray spectrum. Timing measurements (e.g. by the Rossi X-ray Timing satellite RXTE or the X-ray Multi-Mirror Mission XMM-Newton) have produced power density spectra of the fluctuations in the light flux of binaries and showed low and high frequency peaks in the PSD, known as quasi-periodic oscillations (QPOs). They were first noticed in dwarf novae, i.e., erupting cataclysmic variables, as incoherent pulses with time scales of 30-170s along with coherent periodic oscillations with a period of 20s, the so called dwarf nova oscillations (DNOs).

These flux variations occur at all sorts of time scales and some of them are supposedly caused by an ensemble of waves and oscillation modes in the innermost region of the accretion disc. In general, oscillations are the result of restoring forces acting on perturbations. For instance, if one perturbs a fluid element radially inwards, it conserves its own angular momentum and will be rotating too slow for its new location. Centripetal forces consequently push it outwards again. These kind of inertial oscillations are the epicyclic oscillations. The field of studying the temporal behaviour of discs by means of oscillations is called discoseismology

Thermal-viscous instability and the limit cycle

When, at a fixed radius, the \(\dot{M}-\Sigma\) relation for a thin (slim) disc is shaped as an S-curve, with the upper and lower branches being stable, and the middle branch being thermally and viscously unstable, a limit cycle behaviour may occur when the mass supply \(\dot{M}_0\) is in the unstable range. Such a schematic picture, is useful in understanding the basic reason for the disc instability and the limit cycle behaviour, but it can be misleading, because although the instability is local the limit cycle is a global process and quite often the resulting local disc behaviour does not correspond to the simple, schematic S-curve diagram.






Figure 1: Schematic S-curve showing the (local) limit-cycle behaviour in accretion discs. (Figure credits: Kato et al., 1998)   Figure 2: Theoretical two-alpha Disc Instability Model (DIM) to explain the limit-cycle bahaviour seen in cataclysmic variables. (Figure credits: Lasota, 2001)   Figure 3: Scheme of a dwarf nova outburst in cataclysmic variables. (Figure credits: ???)

Cataclysmic variables, low accretion rates: The instability is related to partial hydrogen ionisation, which implies that convection may become the dominant mode of energy transport. The resulting limit cycle is a main ingredient for explaining the dwarf nova outbursts. However, other mutually related, effects also play a role. Particularly important are the varying mass-transfer rate from the secondary, (self-) illumination of the secondary and the disc, tidal instabilities. For the model to work, it is necessary to assume that \(\alpha\) on the upper branch is significantly larger than on the lower branch.

X-ray binaries, high (around Eddington) accretion rates: Theory predicts that the instability should occur for radiation pressure supported discs with luminosities in excess of $L > 0.01L_{Edd}$. Observations argue against the existence of this instability. Stellar-mass black hole sources cross this limit both during their rise to peak luminosity and on their decline to quiescence, showing no symptoms of unstable behaviour. Instead, observations suggest that discs in black hole X-ray binaries are stable up to at least \(L \approx 0.5L_{Edd}\).

Because theory predicts that the instability occurs when the viscous stress is proportional to the total pressure, as in the standard Shakura-Sunyaev alpha-viscosity prescription, other viscosity prescriptions have been suggested (Ciesielski et al. (2012)).

Thin discs

One can then express the Eulerian perturbations of all physical quantities through a single function \(\delta W \propto \delta p/\rho\) which satisfies a second-order partial differential equation. Since the accretion disc is considered to be stationary and axisymmetric, the angular and time dependencies are factored out as \(\delta W = W(r,z)e^{i(m\phi - \sigma t)}\ ,\) where the eigenfrequency \(\sigma(r,z) = \omega - m\Omega\) and \(m\) is the azimuthal wave number. It is assumed that the variation of oscillation modes in radial direction is much stronger than in vertical direction. The resulting two (separated) partial differential equations for the functional amplitude \(W(r,z) = W_r(r) W_y(r,y)\) are given by

\(\frac{d^2W_r}{dr^2}-\frac{1}{(\omega^2-\omega_r^2)}[\frac{d}{dr}(\omega^2-\omega_r^2)]\frac{dW_r}{dr} +\alpha^2(\omega^2-\omega_r^2)(1-\frac{\Psi}{\tilde{\omega}^2})W_r = 0\) and

\((1-y^2)\frac{d^2W_y}{dy^2}-2gy\frac{dW_y}{dy}+2g\tilde{\omega}^2[1-(1-\frac{\Psi}{\tilde{\omega}^2})(1-y^2)]W_y = 0\ .\)

The radial eigenfunction, \(W_r\ ,\) varies fast with \(r\) and the vertical eigenfunction, \(W_z\ ,\) varies slowly with \(r\ .\) The radial and vertical epicyclic frequencies are given by \(\omega_r(r)\) and \(\omega_{\theta}(r)\ ,\) respectively, \(\tilde{\omega}(r) = \omega(r)/\omega_{\theta}(r)\) and \(\alpha(r) = \frac{dt}{d\tau}\frac{\sqrt{g_{rr}}}{c_s(r,0)}\ ,\) using coefficients of the Kerr metric in Boyer-Lindquist coordinates. \(\Psi\) is the eigenvalue of the (WKB) separation function. The radial boundary conditions depend on the type of mode and its capture zone (see below). Oscillations in accretion discs are studied by means of \(\Psi(r,\sigma)\) with the angular mode number \(m\ ,\) the vertical and radial mode numbers (number of nodes in the corresponding eigenfunction) \(j\) and \(n\ ,\) respectively.




Figure 1: Schematic picture showing trapped axisymmetric g-modes with n=1 in the region between \(r_1\) and \(r_2\ ,\) below the radial epicyclic mode \(\kappa(r)=\omega_r(r)\ .\) p-modes only exist above the Keplerian frequency \(\Omega_K\ .\) (Figure credits: Kato et al., 1998)   Figure 2: Schematic picture showing the dependence of three characteristic disc frequencies on the spin \(a\) of the black hole. (Figure credits: Wagoner, 1999)


A mode oscillates in the radial range outside the inner disc, \(r > r_i\ ,\) where

\( (\omega^2 - \omega_r^2) (1 - \frac{\Psi}{\tilde{\omega}^2}) > 0\)

The two points \(r=r_{\pm}(m, a, \sigma)\) refer to the location where the Lindblad resonances occur.

  • p-modes are inertial acoustic modes defined by \(\Psi < \tilde{\omega}^2\) and are trapped where \(\omega^2 > \omega_r^2\) in two zones between the inner (\(r_i\)) and outer (\(r_o\)) radius of the disc. The inner p-modes are trapped between the inner disc edge and the radial epicyclic frequency at \(r_i < r < r_- \) where the gas already is accreted rapidly. The outer p-modes occur at \(r_+ < r < r_o\ .\) The latter, shown in Figure 1, produce the stronger luminosity modulation. In the corotating frame these modes appear at frequencies slightly higher that the radial epicyclic frequency. Pressure is their main restoring force.
  • g-modes are inertial gravity modes defined by \(\Psi > \tilde{\omega}^2\ .\) They are trapped where \(\omega^2 < \omega_r^2\) in the zone \(r_- < r < r_+\) given by the radial dependence of \(\omega_r\ ,\) i.e., g-modes are gravitationally captured in the cavity of the radial epicyclic frequency and are thus the most robust among the modes. Since this is the region of the temperature maximum of the disc, g-modes are expected to be observed best. In the corotating frame these modes appear at low frequencies. Gravity is their main restoring force. - In Figure 1 \(r_1=r_-\) and \(r_2=r_+\) and in figure 2 the spin dependency is plotted for \(m=0\ .\)
  • c-modes are corrugation modes defined by \(\Psi = \tilde{\omega}^2\ .\) They are non-radial (\(m=1\)) and vertically incompressible modes that appear near the inner disc edge and precess slowly around the rotational axis. The c-modes are controlled by the radial dependence of the vertical epicyclic frequency. In the corotating frame these modes appear at highest frequencies.

All modes have frequencies \(\propto 1/M\ .\) Upon the introduction of a small viscosity (\(\nu \propto \alpha, \,\, \alpha << 1\)) most modes grow at a dynamical timescale such that the disc should become unstable.

Thick discs

Geometrically thick discs, i.e. stationary tori, always allow axisymmetric, incompressible modes corresponding to global oscillations of the entire torus at radial (\(\sigma=\omega_r\)) and vertical (\(\sigma=\omega_{\theta}\)) epicyclic frequencies. Other possible modes, provided \(m=0\ ,\) are basically acoustic (p-), surface gravity (g-) and internal inertial (c-) modes and can be found by solving the relativistic Papaloizou-Pringle equation

\( \frac{1}{(-g)^{1/2}}\left\{\partial_{\mu}\left[(-g)^{1/2}g^{\mu\nu}f^n\partial_{\nu}W\right]\right\} - \left(m^2g^{\phi\phi} - 2m\omega g^{t\phi} + \omega^2 g^{tt}\right)f^n W = -\frac{2nA(\bar{\omega}-m\bar{\Omega})^2}{\beta^2 r^2_0}f^{n-1}W \)

together with the boundary condition that the Lagrangian perturbation in pressure at the unperturbed surface (\(f=0\)) vanishes\[ \Delta p = (\delta p + \xi^{\alpha}\nabla_{\alpha}p) = 0 \]


  • \(\times\)-modes are surface gravity modes (k=2) derived from an eigenfunction \(W = a xy\ ,\) for some constant \(a\ ,\) which is odd in \(x\) and \(y\) and results in two modes.

\( \bar{\sigma}_0^2 = \frac{1}{2}\{\omega_r^2+\omega_{\theta}^2 \pm [(\omega_r^2+\omega_{\theta}^2)^2+4\kappa_0^2\omega_{\theta}^2]^{1/2}\}, \)

where \(\bar{\sigma}_0^2=\sigma_0/\Omega_0\) taken at the location \(r_0\) of the pressure maximum in the torus centre (hence the index 0). \(\kappa_0^2 = \frac{\mathcal{E}_0^2}{l_0 A_0^2}(\frac{g^{tt}_{,r}-l_0 g^{t\phi}_{,r}}{g_{rr}}\frac{dl}{dr})\) is the squared frequency of the inertial oscillation in the fluid due to an angular momentum gradient. For constant angular momentum distribution this term vanishes. The positive square root of which gives the x-mode that is a surface gravity mode (Figure 8). The negative square root gives a purely incompressible inertial (c-) mode whose poloidal velocity field represents a circulation around the pressure maximum.

  • breathing- and \(+\)-modes are derived from an eigenfunction \(W = a + bx + cy\ ,\) for some constants \(a,b,c\ .\) The resulting eigenfrequencies are (for \(b=c=0\)) the zero corotation frequency mode as well as

\( \bar{\sigma}_0^2 = \frac{1}{2n}\{(2n+1)(\omega_r^2+\omega_{\theta}^2)-(n+1)\kappa_0^2 \pm [ ((2n+1)(\omega_{\theta}^2-\omega_r^2)^2 + (n+1)\kappa_0^2)^2 + 4(\omega_r^2 - \kappa_0^2) \omega_{\theta}^2 ]^{1/2} \}, \)

    • breathing - modes have frequencies corresponding to the upper sign in the above equation. The torus cross section contracts and expands (Figure 6). Breathing modes are comparable to acoustic modes (k=0, j=1) in the incompressible Newtonian limit for \(l=const.\ ,\) while in the Keplerian limit the mode frequency becomes that of a vertical acoustic wave.
    • \(+\)-modes have frequencies corresponding to the lower sign. In the incompressible \(n\rightarrow 0\) limit they are comparable to (k=2) gravity modes.
Slender r0.gif
Figure 3: Non-oscillating torus
Slender r.gif
Figure 4: Radial epicyclic oscillations
Slender z.gif
Figure 5: Vertical epicyclic oscillations
Slender x.gif
Figure 6\[\times\]-mode oscillations
Slender b.gif
Figure 7: breathing mode oscillations
Slender p.gif
Figure 8\[+\]-mode oscillations

To non-axisymmetric oscillation modes, however, accretion tori are dynamically unstable. The instability, discovered by Papaloizou & Pringle (1984), affects all non-accreting torus configurations, and most violently tori with a constant angular momentum distribution.

Whether or not hydrodynamical oscillation modes may survive such global instabilities or the presence of a weak magnetic field (MRI turbulence), is subject of current, numerical investigations.


Theory versus observations: the electromagnetic spectra

Theoretical models predict both line and continuous electromagnetic spectra of accretion discs. Comparison with observations shows a rather good general agreement. However, several details are not satisfactory fitted and others are fixed by empirical ad hoc assumptions. Generally, the radiative transfer through the disc is not treated as accurately as in the case of stellar atmospheres. The two most acute difficulties that one faces in the accretion disc case are: (a) the disc geometry calls for a 3-D treatment of the problem, (b) the viscous energy generation rate is not known as a function of the position in the disc. Fundamental papers that discuss methods to deal with these difficulties; see e.g. Hubeny et al. (2001) and Davis et al. (2005). In the black hole and neutron star accretion discs, the fact that radiation moves in a highly curved space (spacetime) further complicates the problem of calculating the spectra. Several "ray-tracing" methods of calculating photon trajectories along their geodesics from the source to a distant observer's screen have been developed to deal with this issue; see e.g. Vincent et al. (2011) and Jean-Alain Marck (who was a pioneer of modern ray-tracing methods).

The protoplanetary discs appear in the spectra of young stellar objects as infra-red excess on top of a stellar blackbody signature. Depending on the stage of the system, the excess hump can dominate (class I) the emission profile or is barely notable (class III). Protoplanetary discs contain a large amount of dust, e.g., silicates, minerals or water, whose chemical composition and geometrical shape show up as a forest of line features in the spectrum and greatly influence the opacity and evolution of the disc (Dullemond & Dominik, 2005).

The Cataclysmic Variables (CVs) with accretion discs (i.e., the WD is not or only weakly magnetised) can be of very different types. The bright dwarf novae discs are controlled by a temporary enhancement of the rate of mass transfer. The underluminous and cold dust or debris discs around WD are the result of tidally disrupted asteroids or planetoids. Generally, discs of CVs are divided into two states:

  • the high state (high luminosity, active mass-transfer)
  • the low state (low luminosity, quiescent)

The X-ray Galactic binaries (BHBs) display a number of distinct spectral states (Zdziarski & Gierlinski 2004; Remillard & McClintock, 2006; Done et al., 2007):

  • the very high state (very luminous)
  • the thermal state (disc dominated, luminous, high/soft state),
  • the intermediate state
  • the power-law state (slightly less luminous, low/hard state with strong variability) ,
  • the quiescent state (very underluminous).

The thermal state is well described by the thin disc model and a multi-colour disc (MCD) blackbody model (e.g., diskbb, Mitsuda et al. 1984; Zimmerman et al. 2005; available in XSPEC, Arnaud et al. 1996). Recently, fully relativistic versions of the MCD model for arbitrary BH spin (kerrbb, Li et al. 2005; bhspec, Davis & Hubeny 2006) have been developed, based on the relativistic thin disc model of Novikov & Thorne (1973). These models provide excellent fits to the X-ray spectra of BHBs in the thermal and intermediate state.

Examples of continuous spectra

Shakura-Sunyaev disc spectra


The BeppoSAX observation of LMC X-3 in the thermal (high/soft) state in comparison with the best fit BHSPEC models (for inclination i = 67°, distance D = 52 kpc and viscosity prescription \(\alpha = 0.01\)). The total model is given by the green curve, BHSPEC (red, long-dashed curve) and COMPTT (violet, short-dashed curve) are plotted model components. The unabsorbed BHSPEC model is shown by the orange, solid curve. Figure credits: Davis et al. (2006)

ADAF spectra


The observed low-state spectrum of the X-Ray Nova XTE J1118+480 (crosses, triangels and squares) compared with the two spectra calculated with the ADAF model (red and green lines). Figure taken from Esin et al. (2001)

Slim discs spectra


The soft hump and hard X-ray spectrum of RE J1034+396. The data are taken from Puchnarewicz et al. (2001). The dashed, solid and dotted lines represent slim disc models with \({\dot M}/{\dot M}_{Edd} =\) 5, 10, and 20 respectively. In all cases the central black hole has the mass \(2.25 \times 10^6 M_{sun}\ .\) The fits include Comptonization. Figure taken from Wang & Netzer (2003).

Ion tori spectra


Impact of spin on an ion torus spectrum: a = 0 (solid blue), 0.5M (dashed red) or 0.9M (dotted green). The open circles and the "bow tie" mark observational data of Sgr A*. Note that this is merely a comparison: the spectra do not represent best fits.

Figure taken from Straub et al. (2012)

Examples of line spectra

Müller.camenzind.2004 fe-line-profile.jpg


Vaughan.fabian.2003 fe-line mcg30-6-15.jpg

Figure 1: Schematic profile of the fluorescent iron line. (Figure credits: Müller & Camenzind, 2004)   Figure 2: Observed Fe K alpha line in the Seyfert 1 galaxy MCG 30-6-15. (Figure credits: Vaughan & Fabian, 2003)  

Accretion discs often show all kinds of emission and absorption lines. The most prominent line is the fluorescent emission line of neutral or mildly ionised iron. It occurs in the X-ray band (depending on the degree of ionisation between 6-7 keV), i.e., at energies that can only be created close to a central compact object. The fluorescent Fe line is thus used to probe the vicinity of compact objects. The X-ray line photons are produced when X-rays from the hot inner flow (corona) irradiate the underlying cool, weakly ionised disc. The iron in the disc absorbs the hard X-rays whereby an electron gets excited from the ground state (K-shell) to a higher energy level. After a while it assumes again the ground state by releasing the excess energy as a photon. The strongest line is the Fe \(K_{\alpha}\) line at ~6.4 keV. Due to non-relativistic (Doppler) and special relativistic (beaming) and general relativistic effects (gravitational red-shift) the line is not a simple Gaussian spike but a skewed and asymmetric profile.

Examples of images

The reason why there are so far no direct observations of the immediate environment of black holes is that observed from Earth they have a very small apparent size in the sky. The largest of all is the supermassive black hole in the centre of our galaxy, Sagittarius A*: The apparent diameter of its event horizon is only 53 micro arcseconds. Such dimensions have long been out of reach for astronomical instruments. However, a new generation of interferometers will change that in the coming years (first observations in 2015-2020). The Event Horizon Telescope (EHT), operating in the sub-millimetre radio range will perform Very Large Baseline Interferometry (VLBI) using a network of antennas distributed over the whole globe.

Only those photons of an accretion disc which escape the black hole environment are observable. In an image, a black hole is thus silhouetted dark against the radiant accretion flow. The black hole silhouette itself is confined by a ring of photons that originate from immediately outside the horizon (second order image of the accretion flow) and is calculated by integrating the photon trajectories through spacetime from the observer towards the black hole. This method is called ray-tracing. The image and silhouette that emerge from this calculation and that may be observed by EHT in the near future are rich in information about the properties of the accretion flow and in particular of the black hole.

Figure 12 shows images of a toroidal accretion flow (ion torus) around a supermassive Kerr black hole located at the galactic centre as seen by an observer on Earth. Top left: black hole spin \(a_*=0.5\), torus inclination towards the line of sight inc=80°, torus size \(\lambda=0.3\). Top right: The same as in top left but with \(a_*=0.9\). Bottom left: The same as in top left but with \(\lambda = 0.7\). Bottom right: The same as in top left but with inc=40°.

Figure 12: Images of an ion torus around Sagittarius A* as seen from Earth. Figure credit: Straub et al. (2012)


On line monographs and lecture courses on the accretion discs theory (highly recommended for students)

Original (often cited) discovery papers and influential reviews (available on-line)

  • Zeldovich Ya.B., Novikov, I. D., 1964, Dokl. Akad. Nauk SSSR, 158, 311 (non available on-line)

Monographs and textbooks (non available on-line)

  • Abramowicz M.A., Björnsson G., Pringle J.E., 1999, Theory of Black Hole Accretion Discs, Cambridge University Press, Cambridge
  • Frank J., King A., Raine D.J., 2002, Accretion Power in Astrophysics: Third Edition, Cambridge University Press, Cambridge,
  • Kato S., Fukue J., Mineshige S., 1998, Black-hole accretion disks, Kyoto University Press, Kyoto

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