The motivation for considering higherdimensional black holes with a cosmological constant arises from the AdS/CFT correspondence [1]. This is an equivalence between string theory on spacetimes asymptotic to , where is a compact manifold, and a conformal field theory (CFT) defined on the Einstein universe , which is the conformal boundary of . The best understood example is the case of type IIB string theory on spacetimes asymptotic to , which is dual to superYangMills theory on . Type IIB string theory can be replaced by IIB supergravity in the limit of large and strong ’t Hooft coupling in the YangMills theory.
Most studies of black holes in the AdS/CFT correspondence involve dimensional reduction on to obtain a dimensional gauged supergravity theory with a negative cosmological constant. For example, one can reduce type IIB supergravity on to obtain , gauged supergravity. One then seeks asymptotically antide Sitter blackhole solutions of the gauged supergravity theory. This is certainly easier than trying to find solutions in ten or eleven dimensions. However, one should bear in mind that there may exist asymptotically blackhole solutions that cannot be dimensionally reduced to dimensions. Such solutions would not be discovered using gauged supergravity.
In this section we shall discuss asymptotically blackhole solutions of the gauged supergravity theories arising from the reduction of or supergravity on spheres. The emphasis will be on classical properties of the solutions rather than their implications for CFT. In AdS, linearized supergravity perturbations can be classified as normalizable or nonnormalizable according to how they behave near infinity [1]. By “asymptotically AdS” we mean that we are restricting ourselves to considering solutions that approach a normalizable deformation of global AdS near infinity. A nonnormalizable perturbation would correspond to a deformation of the CFT, for instance, making it nonconformal. Blackhole solutions with such asymptotics have been constructed, but space prevents us from considering them here.
The simplest example of an asymptotically AdS black hole is the SchwarzschildAdS solution [172, 250]:
where is proportional to the mass, and is the radius of curvature of the AdS ground state^{11}. The solution has a regular horizon for any . Definitions of mass and angular momentum for asymptotically AdS spacetimes have been given in [4] and [3]. The mass of SchwarzschildAdS relative to the AdS ground state is [250] For , the stability of SchwarzschildAdS against linearized gravitational perturbations has been proven in [163]. For , spherical symmetry enables one to decompose linearized gravitational perturbations into scalar/vector/tensor types. The equations governing each type can be reduced to ODEs of Schrödinger form, and the stability of vector and tensor perturbations can be established [151]. Stability with respect to scalar gravitational perturbations has not yet been established.It is expected that the SchwarzschildAdS black hole is the unique, static, asymptotically AdS, blackhole solution of vacuum gravity with a negative cosmological constant, but this has not been proven.
The thermodynamics of SchwarzschildAdS were discussed by Hawking and Page for [134] and Witten for [250]. Let denote the horizon radius of the solution. For a small black hole, , the thermodynamic properties are qualitatively similar to those of an asymptoticallyflat Schwarzschild black hole, i.e., the temperature decreases with increasing so the heat capacity of the hole is negative (as is a monotonic function of ). However, there is an intermediate value of at which the temperature reaches a global minimum and then becomes an increasing function of . Hence the heat capacity of large black holes is positive. This implies that the black hole can reach a stable equilibrium with its own radiation (which is confined near the hole by the gravitational potential at large ). Note that for there are two blackhole solutions with the same temperature: a large one with positive specific heat and a small one with negative specific heat.
These properties lead to an interesting phase structure for gravity in AdS [134]. At low temperature, , there is no blackhole solution and the preferred phase is thermal radiation in AdS. At , black holes exists but have greater free energy than thermal radiation. However, there is a critical temperature beyond which the large black hole has lower free energy than thermal radiation and the small black hole. The interpretation is that the canonical ensemble for gravity in AdS exhibits a (firstorder) phase transition at .
In the AdS/CFT context, this Hawking–Page phase transition is interpreted as the gravitational description of a thermal phase transition of the (strongly coupled) CFT on the Einstein universe [249, 250].
When oxidized, to ten or eleven dimensions, the radius of a small SchwarzschildAdS black hole will be much less than the radius of curvature () of the internal space . This suggests that the black hole will suffer from a classical Gregory–Laflammetype instability. The probable endpoint of the instability would be a small black hole localized on , and therefore would not admit a description in gauged supergravity. Since the radius of curvature of is typically and the black hole is much smaller than , the geometry near the hole should be well approximated by the ten or elevendimensional Schwarzschild solution (see e.g., [141]). However, an exact solution of this form is not known.
If we consider pure gravity with a negative cosmological constant then the most general known family of asymptoticallyAdS blackhole solutions is the generalization of the Kerr–Myers–Perry solutions to include a cosmological constant. It seems likely that black rings would exist in asymptoticallyAdS spacetimes, but no exact solutions are known.^{12}
The KerrAdS solution was constructed long ago [27]. It can be parameterized by its mass and angular momentum , which have been calculated (using the definitions of [4]) in [113]. The region of the plane covered by these black holes is shown in Figure 13. Note that, in AdS, angular momentum is a central charge [108]. Hence regular vacuum solutions exhibit a nontrivial lower bound on their mass: . The KerrAdS solution never saturates this bound.

The Myers–PerryAdS solution was obtained in [133] for and for with rotation in a single plane. The general solution was obtained in [111, 112]. They have horizons of spherical topology. There is some confusion in the literature concerning the conserved charges carried by these solutions. A careful discussion can be found in [113]. The solutions are uniquely specified by their mass and angular momenta. For , the region of space covered by the Myers–PerryAdS solution is shown in Figure 14.
Kerr–Myers–PerryAdS solutions have the same symmetries as their asymptoticallyflat cousins, and exhibit similar enhancement of symmetry in special cases. The integrability of the geodesic equation and separability of the Klein–Gordon equation also extends to this case [207, 173, 90].
These solutions reduce to the SchwarzschildAdS solution in the limit of zero angular momentum. It has been shown that the only regular stationary perturbations of the SchwarzschildAdS solution are those that correspond to taking infinitesimal angular momenta in these rotating solutions [162]. Hence, if other stationary vacuum blackhole solutions exist (e.g., black rings) then they are not continuously connected to the SchwarzschildAdS solution.
These solutions exhibit an important qualitative difference from their asymptoticallyflat cousins. Consider the Killing field tangent to the nullgeodesic generators of the horizon:
In asymptoticallyflat spacetime, this Killing field is spacelike far from the black hole, which implies that it is impossible for matter to corotate rigidly with the hole (i.e., to move on orbits of ). However, in AdS, if then is timelike everywhere outside the horizon. This implies that rigid corotation is possible; the Killing field defines a corotating reference frame. Consequently, there exists a Hartle–Hawking state describing thermal equilibrium of the black hole with corotating thermal radiation [133].The dual CFT interpretation is of CFT matter in thermal equilibrium rotating around the Einstein universe [133]. There is an interesting phase structure, generalizing that found for SchwarzschildAdS [133, 13, 20, 135]. For sufficiently large black holes, one can study the dual CFT using a fluid mechanics approximation, which gives quantitative agreement with blackhole thermodynamics [14].
What happens if ? Such black holes are believed to be classically unstable. It was observed in [135] that rotating black holes in AdS may suffer from a superradiant instability, in which energy and angular momentum are extracted from the black hole by superradiant modes. However, it was proven that this cannot occur if . But if then an instability may be present. This makes sense from a dual CFT perspective; configurations with would correspond to CFT matter rotating faster than light in the Einstein universe [133]. The existence of an instability was first demonstrated for small KerrAdS black holes in [23]. A general analysis of odddimensional black holes with equal angular momenta reveals that the threshold of instability is at [177], i.e., precisely where the stability argument of [135] fails. The endpoint of this classical bulk instability is not known.
In , Figure 13 reveals (using ) that all extremal KerrAdS black holes have and are, therefore, expected to be unstable. We have checked that extremal Myers–PerryAdS black holes also have and so they too should be classically unstable. However, the instability should be very slow when the blackhole size is much smaller than the AdS radius , and one expects it to disappear as : it takes an increasingly long time for the superradiant modes to bounce back off the AdS boundary.
Finally, we should mention a subtlety concerning the use of the term “stationary” in asymptotically AdS spacetimes [177]. Consider the metric
This admits several types of globallytimelike Killing fields. For example, there is the “usual” generator of time translations , which has unbounded norm, but there is also the “rotating” Killing field , which has constant norm. On the conformal boundary, is timelike and is null. Hence, from a boundary perspective, particles following orbits of are rotating at the speed of light. These two different types of timelike Killing vector field allow one to define two distinct notions of stationarity for asymptoticallyAdS spacetimes. So far, all known black hole solutions are stationary with respect to both definitions because they admit global Killing fields analogous to , . However, it is conceivable that there exist AdS black holes (with less symmetry than known solutions) that are stationary only with respect to the second definition, i.e., they admit a Killing field that behaves asymptotically like but not one behaving asymptotically like . From a boundary CFT perspective, such black holes would rotate at the speed of light.In order to discuss charged antide Sitter black holes we will need to specify which gauged supergravity theories we are interested in. The bestunderstood examples arise from the dimensional reduction of or dimensional supergravity theories on spheres to give theories with maximal supersymmetry and nonAbelian gauge groups. However, most work on constructing explicit black hole solutions has dealt with consistent truncations of these theories, with reduced supersymmetry, in which the nonAbelian gauge group is replaced by its maximal Abelian subgroup. Indeed, there is no known blackhole solution with a nontrivial nonAbelian gauge field obeying normalizable boundary conditions.
There is a consistent dimensional reduction of supergravity on to give , , gauged supergravity [67]. This nonAbelian theory can be consistently truncated to give , , gauged supergravity, whose bosonic sector is Einstein gravity coupled to four Maxwell fields and three complex scalars [55]. The scalar potential is negative at its global maximum. The ground state of the theory has the scalars taking constant values at this maximum. One can truncate this theory further by taking the scalars to sit at the top of their potential, and setting the Maxwell fields equal to each other. This gives minimal , gauged supergravity, whose bosonic sector is Einstein–Maxwell theory with a cosmological constant. The embedding of minimal , gauged supergravity theories into supergravity can be given explicitly [31], and is much simpler than the embedding of the nonAbelian theory.
The supergravity theory can also be dimensionally reduced on to give , , gauged supergravity [201, 202].
The massive IIA supergravity can be dimensionally reduced on to give gauged supergravity [59]. This theory has halfmaximal supersymmetry.
It is believed that the type IIB supergravity theory can be consistently reduced on to give , , gauged supergravity, although this has been established only for a subsector of the full theory [62]. This theory can be truncated further to give , , gauged supergravity with three vectors and two scalars. Again, setting the scalars to constants and making the vectors equal gives the minimal gauged supergravity, whose bosonic sector is Einstein–Maxwell theory with a negative cosmological constant and a Chern–Simons coupling. The explicit embeddings of these Abelian theories into type IIB supergravity are known [31, 55].
It is sometimes possible to obtain a given lowerdimensional supergravity theory from several different compactifications of a higherdimensional theory. For example, minimal gauged supergravity can be obtained by compactifying type IIB supergravity on any SasakiEinstein space [18]. More generally, if there is a supersymmetric solution of type IIB supergravity that is a warped product of with some compact manifold , then type IIB supergravity can be dimensionally reduced on to give minimal gauged supergravity [104]. An analogous statement holds for compactifications of supergravity to give minimal , gauged supergravity or minimal gauged supergravity [103, 104].
The Reissner–NordströmAdS black hole is a solution of minimal gauged supergravity. It is parameterized by its mass and electric and magnetic charges ,. This solution is stable against linearized perturbations within this (Einstein–Maxwell) theory [164]. Compared with its asymptoticallyflat counterpart, perhaps the most surprising feature of this solution is that it never saturates a BPS bound. If the mass of the black hole is lowered, it will eventually become extremal, but the extremal solution is not BPS. If one imposes the BPS condition on the solution, then one obtains a naked singularity rather than a black hole [221, 185].
Static, sphericallysymmetric, charged, blackhole solutions of the , , gauged supergravity theory were obtained in [69]. The solutions carry only electric charges and are parameterized by their mass and electric charges . Alternatively they can be dualized to give purely magnetic solutions. Once again, they never saturate a BPS bound. One would expect the existence of dyonic solutions of this theory, but such solutions have not been constructed.
Static, sphericallysymmetric, charged, blackhole solutions of , gauged supergravity were obtained in [9]. They are parameterized by their mass and electric charges . If the charges are set equal to each other then one recovers the Reissner–Nordström solution of minimal gauged supergravity. The solutions never saturate a BPS bound.
A static, sphericallysymmetric, charged blackhole solution of , gauged supergravity was given in [59]. Only a single Abelian component of the gauge field is excited, and the solution is parameterized by its charge and mass.
Static, sphericallysymmetric, charged, blackhole solutions of , gauged supergravity are known [55]. They can be embedded into a truncated version of the full theory in which there are two Abelian vectors and two scalars. They are parameterized by their mass and electric charges.
Electricallycharged, asymptoticallyAdS, blackhole solutions exhibit a Hawking–Page like phase transition in the bulk, which entails a corresponding phase transition for the dual CFT at finite temperature in the presence of chemical potentials for the Rcharge. This has been studied in [31, 56, 57, 32].
These black holes exhibit an interesting instability [121, 122]. This is best understood for a black hole so large (compared to the AdS radius) that the curvature of its horizon can be neglected, i.e., it can be approximated by a black brane. The dual CFT interpretation is as a finite temperature configuration in flat space with finite charge density. For certain regions of parameter space, it turns out that the entropy increases if the charge density becomes nonuniform (with the total charge and energy held fixed). Hence, the thermal CFT state exhibits an instability. Using the AdS/CFT dictionary, this maps to a classical instability in the bulk in which the horizon becomes translationally nonuniform, i.e., a Gregory–Laflamme instability. The remarkable feature of this argument is that it reveals that a classical Gregory–Laflamme instability should be present precisely when the black brane becomes locally thermodynamically unstable. Here, local thermodynamic stability means having an entropy, which is concave down as a function of the energy and other conserved charges (if the only conserved charge is the energy, then this is equivalent to positivity of the heat capacity). The Gubser–Mitra (or “correlated stability”) conjecture asserts that this correspondence should apply to any black brane, not just asymptoticallyAdS solutions. See [128] for more discussion of this correspondence.
For finiteradius black holes, the argument is not so clear cut because the dual CFT lives in the Einstein universe rather than flat spacetime, so finite size effects will affect the CFT argument and the Gubser–Mitra conjecture does not apply. Nevertheless, it should be a good approximation for sufficiently large black holes and hence there will be a certain range of parameters for which large charged black holes are classically unstable.^{13}
The most general, known, stationary, blackhole solution of minimal , gauged supergravity is the KerrNewmanAdS solution, which is uniquely parameterized by its mass , angular momentum and electric and magnetic charges . The thermodynamic properties of this solution, and implications for the dual CFT were investigated in [20]. An important property of this solution is that it can preserve some supersymmetry. This occurs for a oneparameter subfamily specified by the electric charge: , , [171, 21]. Hence supersymmetric black holes can exist in AdS but they exhibit an important qualitative difference from the asymptotically flat case; they must rotate.
Charged rotating blackhole solutions of more general gauged supergravity theories, e.g., , gauged supergravity, should also exist. Electrically charged, rotating solutions of the theory, with the four charges set pairwise equal, were constructed in [36].
Charged, rotating blackhole solutions of , gauged supergravity have been constructed by truncating to a theory [40, 43]. In this theory, one expects the existence of a topologicallyspherical blackhole solution parameterized by its mass, three angular momenta, and two electric charges. This general solution is not yet known. However, solutions with three equal angular momenta but unequal charges have been constructed [40], as have solutions with equal charges but unequal angular momenta [43]. Both types of solution admit BPS limits.
Charged, rotating blackhole solutions of gauged supergravity have not yet been constructed.
The construction of charged rotating blackhole solutions of gauged supergravity has attracted more attention [125, 124, 60, 61, 37, 38, 178, 39, 191]. The most general known blackhole solution of the minimal theory is that of [38]. This solution is parameterized by the conserved charges of the theory, i.e., the mass , electric charge and two angular momenta , . Intuition based on results proved for asymptoticallyflat solutions suggests that, for this theory, this is the most general topologicallyspherical stationary black hole with two rotational symmetries. In the BPS limit, these solutions reduce to a twoparameter family of supersymmetric black holes. In other words, one loses two parameters in the BPS limit (just as for nonstatic asymptoticallyflat black holes in , e.g., the BMPV black hole or BPS black rings).
Analogous solutions of , gauged supergravity are expected to be parameterized by the six conserved quantities , , , , , . However, a sixparameter solution is not yet known. The most general known solutions are the fourparameter BPS solution of [178], and the fiveparameter nonextremal solution of [191], which has two of the charges equal. The former is expected to be the general BPS limit of the yet to be discovered sixparameter blackhole solution (as one expects to lose two parameters in the BPS limit). The latter solution should be obtained from the general sixparameter solution by setting two of the charges equal.
Supersymmetric AdS black holes have , which implies that they rotate at the speed of light with respect to the conformal boundary [125]. More precisely, the corotating Killing field becomes null on the conformal boundary. Hence, the CFT interpretation of these black holes involves matter rotating at the speed of light in the Einstein universe. The main motivation for studying supersymmetric AdS black holes is the expectation that it should be possible to perform a microscopic CFT calculation of their entropy. The idea is to count states in weakly coupled CFT and then extrapolate to strong coupling. In doing this, one has to count only states in short BPS multiplets that do not combine into long multiplets as the coupling is increased. One way of trying to do this is to work with an index that receives vanishing contributions from states in multiplets that can combine into long multiplets. Unfortunately, such indices do not give agreement with blackhole entropy [160]. This is not a contradiction; although certain multiplets may have the right quantum numbers to combine into a long multiplet, the dynamics of the theory may prevent this from occurring, so the index undercounts BPS states.
The fact that these supersymmetric black holes have only four independent parameters is puzzling from the CFT perspective, since BPS states in the CFT carry five independent charges. Maybe there are more general blackhole solutions. It seems unlikely that one could generalize the solutions of [178] to include an extra parameter since then one would also have an extra parameter, in the corresponding nonBPS solutions, which would therefore form a seven parameter family in a theory with only six conserved charges. This seems unlikely for topologicallyspherical black holes. But we know that black rings can carry nonconserved charges, so maybe this points to the existence of supersymmetric AdS black rings. However, it has been shown that such solutions do not exist in minimal gauged supergravity [179]. The proof involves classifying supersymmetric nearhorizon geometries (with two rotational symmetries), and showing that topology horizons always suffer from a conical singularity, except in the limit in which the cosmological constant vanishes. Analogous results for the theory have also been obtained [176]. So if AdS black rings exist then they cannot be “balanced” in the BPS limit.
Maybe the resolution of the puzzle involves 10d black holes with no 5D interpretation, or 5D black holes involving nonAbelian gauge fields, or 5D black holes with only one rotational symmetry. Alternatively, perhaps we already know all the BPS blackhole solutions and the puzzle arises from a lack of understanding of the CFT. For example, maybe, at strong coupling, only a four charge subspace of BPS CFT states carries enough entropy to correspond to a macroscopic black hole.
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