The Einstein–Hilbert action is generalized to higher dimensions in the form

This is a straightforward generalization, and the only aspect that deserves some attention is the implicit definition of Newton’s constant in dimensions. It enters the Einstein equations in the conventional form where . This definition of the gravitational coupling constant, without any additional dimension-dependent factors, has the notable advantage that the Bekenstein-Hawking entropy formula takes the same form in every dimension. This follows, e.g., from the standard Euclidean quantum gravity calculation of the entropy.Mass, angular momenta, and other conserved charges of isolated systems are defined through comparison to the field created near asymptotic infinity by a weakly gravitating system ([154] gives a careful Hamiltonian analysis of conserved charges in higher-dimensional asymptotically-flat spacetimes). The Einstein equations for a small perturbation around flat Minkowski space

in linearized approximation take the conventional form where and we have imposed the transverse gauge condition .Since the sources are localized and we work at linearized perturbation order, the fields in the asymptotic region are the same as those created by point-like sources of mass and angular momentum with antisymmetric matrix , at the origin of flat space in Cartesian coordinates,

The equations are easily integrated, assuming stationarity, to find where , and is the area of a unit -sphere. From here we recover the metric perturbation as It is often convenient to have the off-diagonal rotation components of the metric in a different form. By making a suitable coordinate rotation the angular momentum matrix can be put into block-diagonal form, each block being a antisymmetric matrix with parameter Here labels the different independent rotation planes. If we introduce polar coordinates on each of the planes then (no sum over ) In the last expression we have introduced the ‘direction cosines’Given the abundance of black-hole solutions in higher dimensions, one is interested in comparing properties, such as the horizon area , of different solutions characterized by the same set of parameters . A meaningful comparison between dimensionful magnitudes requires the introduction of a common scale, so the comparison is made between dimensionless magnitudes obtained by factoring out this scale. Since classical general relativity in vacuum is scale invariant, the common scale must be one of the physical parameters of the solutions, and a natural choice is the mass. Thus we introduce dimensionless quantities for the spins and the area ,

where the numerical constants are (these definitions follow the choices in [80]). Studying the entropy, or the area , as a function of for fixed mass is equivalent to finding the function .Note that, with our definition of the gravitational constant , both the Newtonian gravitational potential energy,

and the force law (per unit mass) acquire -dependent numerical prefactors. Had we chosen to define Newton’s constant so as to absorb these factors in the expressions for or , Equation (7) would have been more complicated.To warm up before dealing with black holes, we follow John Michell and Simon de Laplace and compute, using Newtonian mechanics, the radius at which the escape velocity of a test particle in this field reaches the speed of light. The kinetic energy of a particle of unit mass with velocity is , so the equation that determines the Michell–Laplace ‘horizon’ radius is

We will see in the next section 3.2 that, just like in four dimensions, this is precisely equal to the horizon radius for a static black hole in higher dimensions.

Consider the linearized solution above for a static source (14) in spherical coordinates, and pass to a gauge where is the area radius,

The linearized approximation to the field of a static source is then where, to lighten the notation, we have introduced the ‘mass parameter’ This suggests that the Schwarzschild solution generalizes to higher dimensions in the form In essence, all we have done is change the radial falloff of the Newtonian potential to the -dimensional one, . As Tangherlini found in 1963 [232], this turns out to give the correct solution: it is straightforward to check that this metric is indeed Ricci flat. It is apparent that there is an event horizon at , which coincides with the Michell–Laplace result (25).Having this elementary class of black-hole solutions, it is easy to construct other vacuum solutions with event horizons in . The direct product of two Ricci-flat manifolds is itself a Ricci-flat manifold. So, given any vacuum black-hole solution of the Einstein equations in dimensions, the metric

describes a black -brane, in which the black-hole horizon is extended to a horizon , or if we identify periodically . A simple way of obtaining another kind of vacuum solution is the following: unwrap one of the directions , perform a boost , , and re-identify points periodically along the new coordinate . Although locally equivalent to the static black brane, the new boosted black-brane solution is globally different from it.These black brane spacetimes are not (globally) asymptotically flat, so we only introduce them insofar as they are relevant for understanding the physics of asymptotically-flat black holes.

The stability of the Schwarzschild solution against linearized gravitational perturbations can be analyzed by decomposing such perturbations into scalar, vector and tensor types according to how they transform under the rotational-symmetry group [105, 163, 151]. Assuming a time dependence and expanding in spherical harmonics on , the equations governing each type of perturbation reduce to a single ODE governing the radial dependence. This equation can be written in the form of a time-independent Schrödinger equation with “energy” eigenvalue .

In investigating stability, we consider perturbations that are regular on the future horizon and outgoing at infinity. An instability would correspond to a mode with . For such modes, the boundary conditions at the horizon and infinity imply that the left-hand side (LHS) of the Schrödinger equation is self-adjoint, and hence is real. Therefore, an unstable mode must have negative imaginary . For tensor modes, the potential in the Schrödinger equation is manifestly positive, hence and there is no instability [105]. For vectors and scalars, the potential is not everywhere positive. Nevertheless, it can be shown that the operator appearing on the LHS of the Schrödinger equation is positive, hence and there is no instability [151]. In conclusion, the Schwarzschild solution is stable against linearized gravitational perturbations.

The instabilities of black strings and black branes [118, 119] have been reviewed in [167, 128], so we shall be brief in this section and only mention the features that are most relevant to our subject. We shall only discuss neutral black holes and black branes; when charges are present, the problem becomes more complex.

This instability is the prototype for situations in which the size of the horizon is much larger in some
directions than in others. Consider, as a simple, extreme case of this, the black string obtained by adding a
flat direction to the Schwarzschild solution. One can decompose linearized gravitational perturbations
into scalar, vector and tensor types according to how they transform with respect to transformations of the
Schwarzschild coordinates. Scalar and vector perturbations of this solution are stable [117]. Tensor
perturbations that are homogeneous along the -direction are also stable, since they are the same as
tensor perturbations of the Schwarzschild black hole. However, there appears to be an instability for
long-wavelength tensor perturbations with nontrivial dependence on ; the frequency of
perturbations acquires a positive imaginary part when , where is the
Schwarzschild horizon radius. Thus, if the string is compactified on a circle of length , it
becomes unstable. Of the unstable modes, the fastest one (with the largest imaginary frequency) occurs for
roughly one half of . The instability creates inhomogeneities along the direction of the string.
Their evolution beyond the linear approximation has been followed numerically in [42]. It is unclear yet
what the endpoint is; the inhomogeneities may well grow until a sphere pinches down to a
singularity.^{3}
In this case, the Planck scale will be reached along the evolution, and fragmentation of the black string into
black holes, with a consistent increase in the total horizon entropy, may occur.

Another important feature of this phenomenon is the appearance of a zero-mode (i.e., static) perturbation with . Perturbing the black string with this mode yields a new static solution with inhomogeneities along the string direction [117, 120]. Following numerically these static perturbations beyond the linear approximation has given a new class of inhomogeneous black strings [248].

These results easily generalize to black -branes; for a wavevector along the directions tangent to the brane, the perturbations with are unstable. The value of depends on the codimension of the black brane, but not on .

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