Solving higher curvature gravity theories
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Abstract
Solving field equations in the context of higher curvature gravity theories is a formidable task. However, in many situations, e.g., in the context of f(R) theories, the higher curvature gravity action can be written as an Einstein–Hilbert action plus a scalar field action. We show that not only the action but the field equations derived from the action are also equivalent, provided the spacetime is regular. We also demonstrate that such an equivalence continues to hold even when the gravitational field equations are projected on a lowerdimensional hypersurface. We have further addressed explicit examples in which the solutions for Einstein–Hilbert and a scalar field system lead to solutions of the equivalent higher curvature theory. The same, but on the lowerdimensional hypersurface, has been illustrated in the reverse order as well. We conclude with a brief discussion on this technique of solving higher curvature field equations.
Keywords
Scalar Field Field Equation Extra Dimension Conformal Transformation Planck Scale1 Introduction
The energy scales in particle physics are arranged in a hierarchical manner. While the scale of the weak interaction corresponds to \(E\sim 10^{3}~\text {GeV}\), the strong interaction at a scale of \(E\sim 10^{16}~\text {GeV}\) exceeds the weak scale by a factor of \(10^{13}\). This large difference leads to a fine tuning problem in the scheme of renormalization – known as the gauge hierarchy problem. This fine tuning is absolutely necessary to renormalize the mass of the Higgs boson, which recently has been detected with a mass of \(127~\text {GeV}\). At face value, this fine tuning is what nature prefers and the question is why? Hence it is natural to ask, is there a more fundamental principle from which this fine tuning would appear naturally?
There has been a large amount of work to address the hierarchy problem, and a few candidates have emerged out of it – supersymmetry, technicolor, and extra dimensions. In this work we will be concerned solely with the third alternative, i.e., we will assume the actual spacetime has more than three spatial dimensions (commonly referred to as the bulk), while the spacetime we live in is a fourdimensional hypersurface in the bulk (commonly referred to as the brane). The two immediate observable consequences are a change of \(1/r^{2}\) gravitational force law at the length scale of extra dimensions, and the existence of a massive graviton through the Kaluza–Klein tower [1, 2, 3, 4, 5, 6, 7, 8, 9].
To probe these extra dimensions one needs to have a high enough energy or a high enough curvature, such that the relevant energy scale of the problem comes close to the Planck scale. General Relativity, described by the Einstein–Hilbert action is considered to be an effective theory of gravity, valid far below the fundamental Planck scale [10]. Once energies approach the Planck scale, one not only expects to observe deviations from the Einstein–Hilbert action but also signatures of the extra dimensions. This is particularly relevant, since future colliders will probe higher and higher energies such that aspects beyond general relativity should become apparent. Since the ultraviolet behavior of the true gravity theory is yet unknown one hopes that in these high energy/high curvature regimes deviations from standard model or deviations from Einstein gravity may appear through the existence of extra dimensions. To capture some of the aspects of “quantum gravity” one is tempted to consider how the presence of higher curvature (and higher derivative) invariants in the higherdimensional gravitational action modifies the wellknown results [11, 12, 13, 14, 15].
The higher derivative terms that one can add to the Einstein–Hilbert action are not unique. However, many of these terms can lead to a linear instability, called the Ostrogradski instability, leading to the appearance of ghost fields and hence will not be considered in this work. Among the higher curvature theories, the Lanczos–Lovelock gravity and the f(R) gravity are of much importance. The Lanczos–Lovelock gravity is special in the sense that the field equations derived from the Lanczos–Lovelock action contain only second order derivatives of the metric and have a natural thermodynamic interpretation [16, 17, 18, 19, 20]. On the other hand, f(R) gravity was first introduced to explain both early and late time exponential expansion of the universe without invoking additional matter components, e.g., dark energy [21, 22, 23, 24, 25, 26, 27]. But only addressing cosmological observations does not lead to a viable model, for that the f(R) theories should pass the local gravity tests – perihelion precession of Mercury and bending angle of light – as well. It turns out that solar system experiments do not exclude the viability of f(R) theories at scales shorter than the cosmological ones, but they provide constraints on \(f''(R)\) and hence constraints on the parameters of the model. Thus it is can be affirmed that extended gravity theories cannot be ruled out, definitively, using Solar System experiments [28, 29, 30, 31].
It is also well known that f(R) gravity theories can be related to scalar–tensor theories by a conformal transformation at the action level [22, 23, 24, 32, 33, 34, 35, 36, 37, 38]. Thus it is important to consider the following situation – obtaining field equations from the scalar–tensor representation and from the f(R) gravity representation. Since the two actions are related by a conformal transformation, the field equations should also be equivalent. However, the situation is not trivial, since the metric in scalar–tensor representation depends on the conformal factor, its variation can potentially lead to various additional terms, which must cancel other terms exactly in order to arrive at the equivalence. If the equivalence exists, we can use it to solve field equations for scalar–tensor theory and obtain the solution corresponding to f(R) action and vice versa. This would be advantageous, since in general solving the field equations for f(R) gravity, where R is not a constant, is difficult^{1} [22, 41, 42, 43, 44, 45, 46, 47]. The corresponding scalar–tensor solution could in principle be much simpler. The same should work on the brane hypersurface as well. The effective field equations on the brane derived through the Gauss–Codazzi formalism in the f(R) representation [48, 49, 50, 51, 52, 53] should be equivalent to the same but derived from the scalar–tensor representation. The nontriviality of this result originates from the quadratic combination of energy momentum tensor and extrinsic curvature appearing in the effective field equations.
As an aside, we should mention that the conformal transformation is well motivated only when the spacetime does not have singularities. Such singularity free spacetimes have been obtained earlier in the context of cosmology with moduli dependent loop corrections of the gravitational part of superstring effective action with orbifold compactifications [54]. However, to obtain a singularity free description it was necessary that the stress energy tensor associated with the modulus should violate the strong energy condition. In circumstances where the energy conditions are obeyed, one obtains singular solutions in general. For singular spacetimes, viz., cosmological spacetimes near the Big Rip or Big Crunch the transformation can break down. In those contexts it can exhibit peculiar behavior, e.g., the Big Rip singularity, which may appear in some versions of f(R) gravity can either map itself to the infinite past or future, or it can be replaced by a Big Crunch singularity [35, 55]. Another point requires clarification at this stage; this has to do with the physical nonequivalence of the two frames. All the comments phrased above have to do with mathematical equivalence, but the physical solutions can be very different [55]. This is evident, since the conformal factor can change the complete structure of the spacetime. This fact was pointed out earlier in [56] by showing that through a conformal transformation one can create matter and as a result, one frame is empty while another has matter, and clearly they are physically nonequivalent. This should not come as a surprise, since the Schwarzschild metric under conformal transformation no longer satisfies Einstein’s equations. Further in the cosmological context for f(R) gravity model it was explicitly demonstrated [55, 57] that neither the Hubble parameter nor the deceleration parameter matches in Jordan and Einstein frame, showing the physical nonequivalence. In view of the above, the phrase “equivalence” in the following sections should be understood in a mathematical sense, not in a physical sense. Further, we will content ourselves, with only those spacetimes (or regions of spacetimes) which are regular, such that the conformal transformation between the two frames is well defined throughout the region of interest.
The paper is organized as follows: In Sect. 2 we present a brief review of the equivalence between the f(R) gravity and scalar–tensor theory in five dimensions and hence the equivalence between the bulk field equations as well. Section 3 is devoted to show the equivalence between the effective field equations on the brane. The application of the bulk equivalence is presented in Sect. 5. There we have started from scalar–tensor theory and have solved the bulk equations, from which the solution in f(R) representation is obtained. In Sect. 6 we consider brane spacetime where the solution in scalar–tensor representation starting from the f(R) representation is derived as another explicit example. We conclude with a brief discussion of this technique.
We have set the fundamental constants c and \(\hbar \) to unity and shall work with mostly the positive signature of the metric. The Latin indices, \(a,b,\ldots \) runs over the full spacetime indices, while Greek indices, \(\mu ,\nu ,\ldots \) stand for fourdimensional spacetime.
2 Equivalence of gravitational field equations in the bulk
Even though we have used the metric formalism to arrive at the equivalence, one can also use another method known as the Palatini method. For discussions of the same we refer the reader to [23, 58, 59, 60, 61].
3 Equivalence of effective field equations on the brane
In the previous section, we have shown the equivalence between the bulk field equations derived from the Einstein and the Jordan frames. However, from the perspective of brane world, governed by effective field equations derived from the bulk action, the equivalence of the effective field equations is more important. The effective field equations involve various quadratic combinations of the extrinsic curvature and the matter energy momentum tensor. Thus all the additional terms with their appropriate factors present in the Einstein frame must cancel each other such that effective field equations in the Jordan frame are obtained. In this connection we would like to highlight that in most of the works related to f(R) gravity the surface term in the Einstein frame is ignored; however, in order to prove the equivalence on the brane this term is absolutely necessary. Hence the equivalence of the effective field equations too, is a nontrivial statement. In this section we will explicitly demonstrate this.
4 A comparison with reconstruction methods in f(R) gravity
In the above two sections we have shown the equivalence of gravitational field equations both in the bulk and in the brane, respectively. In this section we will present a comparison of our method with an existing wellknown method in f(R) gravity, the reconstruction method. As already emphasized, due to the presence of higher derivatives in the field equations for f(R) gravity, obtaining a straightforward solution in a general case is very difficult. Even for systems with a large number of symmetries, e.g., in cosmology which has a single unknown function a(t), solving the field equations directly in the Jordan frame is very complicated. This leads to the reconstruction method, which we will briefly summarize [63, 64, 65, 66].

An important limitation of the reconstruction method is that only very simple cosmic histories, e.g. simple power law behaviors, can be connected to f(R) theory in an exact way. Our method has similar disadvantages. Even though the field equations can be exactly solved in the Einstein frame for a few cases of interest, the inversion of the potential to f(R) theory can be performed only in simple situations.

The reconstruction method is adapted to cosmological spacetimes only, since the cosmic history is known through experiments. However, the situation we are interested in corresponds to higherdimensional physics in the presence of higher curvature gravity, the possible behavior of the warp factor, and the brane separation. Since there is no experimental backdrop for extra dimensions it is not possible to come up with a physical ansatz. Thus one needs to solve the field equations at face value, which can be efficiently done using our method as we have illustrated in the next sections.

The utility of the reconstruction scheme lies in its quick and straightforward analysis. Given a phenomenological scale factor a(t) one needs to solve a single differential equation to get f(R), given the inversion \(t=t(R)\). In our method one first need to solve the gravity plus scalar field system to get the solution in Einstein frame, which itself is a formidable task. Then one needs to invert the potential \(V(\phi )\) to get f(R) and finally the conformal transformation will yield the solution in the Jordan frame.
5 Einstein to Jordan frame in the bulk: explicit examples
We will now illustrate through simple examples how one might obtain solutions to bulk field equations in f(R) gravity, which involve higher derivative terms, by exploiting the equivalence with scalar tensor theory depicted in the previous sections. As emphasized before, due to the occurrence of higher derivative terms it is difficult to solve for the bulk equations of f(R) gravity. On the other hand, solving a set of coupled equations of gravity plus scalar field system is much simpler. Hence, through the equivalence shown earlier, if we can obtain a solution for the bulk metric \(g_{ab}\) in the Einstein frame, the corresponding solution in the Jordan frame will differ only by a conformal factor.
Before we jump into detailed calculations it is worthwhile to sketch the flowchart we are going to follow – (a) We will start with the bulk action in the Einstein frame. (b) For some suitable potential, we will find the metric describing the bulk spacetime, by solving the bulk field equations. (c) We will match the potential in the Einstein frame with the corresponding f(R) theory in the Jordan frame. (d) Finally the conformal transformation will yield the corresponding bulk metric in the Jordan frame. In the examples to follow we will explicitly illustrate all the four steps mentioned above.
 The superpotential is linear in \(\phi \), i.e., \(W(\phi )=c\phi \). Then from Eq. (32) one obtains \(A'=(b/3)\phi \), such that \(A''=(c/3)\phi '\), while \(\phi '=c/\kappa _{5}^{2}\). This set identically satisfies Eq. (29). The corresponding potential \(V(\phi )\) turns out to bewith the following solution for A(y) and \(\phi (y)\):$$\begin{aligned} V(\phi )=\frac{\Lambda }{\kappa _{5}^{2}}+\frac{c^{2}}{2\kappa _{5}^{4}r_{c}^{2}}\frac{2c^{2}}{3\kappa _{5}^{2}r_{c}^{2}}\phi ^{2}, \end{aligned}$$(33)$$\begin{aligned} \phi (y)&=\phi _{0}+\frac{c}{\kappa _{5}^{2}}y~, \end{aligned}$$(34)$$\begin{aligned} A(y)&=A_{0}+\frac{c\phi _{0}}{3}y+\frac{c^{2}}{6\kappa _{5}^{2}}y^{2}~. \end{aligned}$$(35)
 The superpotential is quadratic in \(\phi \), i.e., \(W(\phi )=ab\phi ^{2}\). From Eq. (32) we get \(A'=(1/3)(ab\phi ^{2})\), hence this yields \(A''=(2b/3)\phi \phi '\), with \(\phi '=(2b/\kappa _{5}^{2})\phi \). These expressions can easily be manipulated to show that Eq. (29) is indeed satisfied. Then the potential becomeswith the following solutions for A(y) and \(\phi (y)\):$$\begin{aligned} V(\phi )&=\left( \frac{\Lambda }{\kappa _{5}^{2}}\frac{2a^{2}}{3r_{c}^{2}\kappa _{5}^{2}}\right) \nonumber \\&\quad +\left( \frac{b^{2}}{2r_{c}^{2}\kappa _{5}^{4}}+\frac{4ab}{3r_{c}^{2}\kappa _{5}^{2}}\right) \phi ^{2}\frac{2b^{2}}{3r_{c}^{2}\kappa _{5}^{2}}\phi ^{4}, \end{aligned}$$(36)$$\begin{aligned} \phi (y)&=\phi _{0}\exp \left( \frac{b}{\kappa _{5}^{2}}y\right) ~,\end{aligned}$$(37)$$\begin{aligned} A(y)&=A_{0}+\sqrt{\frac{\Lambda r_{c}^{2}}{6}}y+\frac{\kappa _{5}^{2}}{6}\phi _{0}^{2}\exp \left( 2\frac{b}{\kappa _{5}^{2}}y\right) . \end{aligned}$$(38)
 The simplest model is always the best to start with. For f(R) gravity this corresponds to a situation, where the Einstein–Hilbert term receives a quadratic correction,^{2} i.e., \(f(R)=R+\alpha R^{2}\). For this particular model of f(R) gravity, the scalar field and the potential can be written in terms of the Ricci scalar in the Jordan frame as (using Eq. (39))$$\begin{aligned} \kappa _{5}\phi&=\frac{2}{\sqrt{3}}\ln ~\left( 1+2\alpha R\right) ;\quad R=\frac{1}{2\alpha }\left[ \exp \left( {\frac{\sqrt{3}}{2}\kappa _{5}\phi }\right) 1\right] ~, \end{aligned}$$(40)The minimum of the potential corresponds to \(\partial V/\partial \phi =0\), with \(\partial ^{2}V/\partial \phi ^{2}>0\). Both conditions can be satisfied, provided \(e^{\kappa _{5}\phi }=1\), or \(\phi =0\). Finally, expanding around the minimum one obtains the following form for the potential:$$\begin{aligned} V(\phi )&=\frac{1}{8\alpha \kappa _{5}^{2}}\left[ \exp \left( \frac{\kappa _{5}\phi }{2\sqrt{3}}\right) 2\exp \left( \frac{\kappa _{5}\phi }{\sqrt{3}} \right) \right. \nonumber \\&\left. \quad \,+\exp \left( \frac{5}{2\sqrt{3}}\kappa _{5}\phi \right) \right] . \end{aligned}$$(41)Thus we have a quadratic potential for \(\phi \), which originates from \(R+\alpha R^{2}\) gravity. Matching the potential to that derived in the Einstein frame, given by Eq. (33), we obtain the following relation: \(\Lambda =9/(128\alpha )\).$$\begin{aligned} V(\phi )&=\frac{1}{8\alpha \kappa _{5}^{2}}\left[ \left( 1+\frac{\kappa _{5}\phi }{2\sqrt{3}}+\frac{1}{2}\frac{\kappa _{5}^{2}\phi ^{2}}{12} \right) 2\left( 1\frac{\kappa _{5}\phi }{\sqrt{3}}+\frac{1}{6}\kappa _{5}^{2}\phi ^{2}\right) \nonumber \right. \\&\left. \qquad +\left( 1\frac{5}{2\sqrt{3}}\kappa _{5}\phi +\frac{1}{2}\frac{25\kappa _{5}^{2}\phi ^{2}}{12}\right) \right] =\frac{3}{32\alpha }\phi ^{2}. \end{aligned}$$(42)
 Let us now consider a more general f(R) gravity model for which \(f(R)=R+\alpha R^{2}+\beta R^{4}\), where \(\alpha \) and \(\beta \) are dimensionful constant coefficients, with values such that the model becomes ghost free. Then from Eq. (39) we obtain \(R=(\sqrt{3}\kappa _{5}\phi )/(4\alpha )\), such that the potential turns out to bewhere we have assumed \(\alpha \gg \beta \gg \alpha ^{2}\), consistent with the ghost free criterion for this f(R) model. Comparing this with the potential obtained by solving bulk field equations in the Einstein frame, presented in Eq. (36) we immediately obtain$$\begin{aligned} V(\phi )=\frac{\alpha R^{2}+3\beta R^{4}}{2\kappa _{5}^{2}}=\frac{3}{32\alpha }\phi ^{2}+\frac{1}{2}\frac{3^{3}\beta \kappa _{5}^{2}}{4^{4}\alpha ^{4}}\phi ^{4}, \end{aligned}$$(43)$$\begin{aligned} a&=\sqrt{\frac{3r_{c}^{2}\Lambda }{2}};\quad b=\frac{4a\kappa _{5}^{2}}{3}\pm \sqrt{\frac{16a^{2}\kappa _{5}^{4}}{9}+\frac{3r_{c}^{2}\kappa _{5}^{4}}{16\alpha }};\nonumber \\ \beta&=\frac{4^{5}b^{2}\alpha ^{4}}{3^{4}r_{c}^{2}\kappa _{5}^{4}}. \end{aligned}$$(44)
 For \(f(R)=R+\alpha R^{2}\), the corresponding scalar field potential in the Einstein frame is quadratic with the mapping being given by Eq. (42). Thus, the conformal factor turns out to be \(\Omega =(1+2\alpha R)^{1/3}=[1+(\sqrt{3}\kappa _{5}\phi /2)]^{1/3}\). Hence the bulk solution in the Jordan frame corresponds towhere c and \(\phi _{0}\) are arbitrary constants of integration. Thus for the quadratic f(R) model under consideration, one can map it to the Einstein frame and obtain the respective potential. For this particular case, the field equations in the Einstein frame become exactly solvable and hence by a conformal transformation one can obtain the corresponding solution in the Jordan frame. Further, from Eq. (45) it turns out that the warp factor is governed by the factors \(c\phi _{0}\) and \(c/\kappa _{5}\). Hence in order to have proper suppression of the Planck scale on the visible brane one must have the conditions \(c<\kappa _{5}\) and \(c\phi _{0}\sim 36\). Hence one arrives at \(\phi _{0}>\kappa _{5}^{1}\). Further, in this model the radion field varies with extra dimension y as \((a+by)^{2/3}\), where a and b depend on \(c,\kappa _{5}\) and \(\phi _{0}\).$$\begin{aligned}&\mathrm{d}s^{2}=\left[ 1+\frac{\sqrt{3}\kappa _{5}\phi (y)}{2}\right] ^{2/3}\nonumber \\&\quad \quad \; \times \left\{ e^{2A(y)}\eta _{\mu \nu }\mathrm{d}x^{\mu }\mathrm{d}x^{\nu }+r_{c}^{2}\mathrm{d}y^{2} \right\} , \\&\phi (y)=\phi _{0}+\frac{c}{\kappa _{5}^{2}}y;\quad A(y)=A_{0}+\frac{c\phi _{0}}{3}y+\frac{c^{2}}{6\kappa _{5}^{2}}y^{2}, \nonumber \end{aligned}$$(45)
 For the other model, i.e., \(f(R)=R+\alpha R^{2}+\beta R^{4}\), the scalar field potential in the Einstein frame is quartic. From this one can relate the parameters \(\alpha ,\beta \) with the respective ones in the Einstein frame. Following the same strategy as above, the conformal factor turns out to yield \(\Omega =(1+2\alpha R+4\beta R^{3})^{1/3}=[1+(\sqrt{3}\kappa _{5}\phi /2)+(3\sqrt{3}\beta \kappa _{5}^{3}\phi ^{3}/16\alpha ^{3})]^{1/3}\). Using this the solution in the Jordan frame becomesThis demonstrates another f(R) theory for which the corresponding potential in the Einstein frame leads to an exact solution. Using this and the mapping between Einstein and Jordan frame one obtains the respective solution in the Jordan frame. However, in contrast with the previous situation, in this case the warp factor behaves exactly like the Randall–Sundrum scenario, since all the corrections are exponentially suppressed (see Eq. (46)). The radion field is almost constant due to identical exponential suppression. Hence the f(R) model with quartic correction is more favored in the extradimensional physics than the earlier one.$$\begin{aligned} \mathrm{d}s^{2}&=\left[ 1+\frac{\sqrt{3}\kappa _{5}\phi (y)}{2}+\frac{3\sqrt{3}\beta \kappa _{5}^{3}\phi (y)^{3}}{16\alpha ^{3}}\right] ^{2/3}\nonumber \\&\quad \times \left\{ e^{2A(y)}\eta _{\mu \nu }\mathrm{d}x^{\mu }\mathrm{d}x^{\nu }+r_{c}^{2}\mathrm{d}y^{2} \right\} , \\ \phi (y)&=\phi _{0}\exp \left( \frac{b}{\kappa _{5}^{2}}y\right) ;\nonumber \\ A(y)&=A_{0}+\sqrt{\frac{\Lambda r_{c}^{2}}{6}}y+\frac{\kappa _{5}^{2}}{6}\phi _{0}^{2}\exp \left( 2\frac{b}{\kappa _{5}^{2}}y\right) ~. \nonumber \end{aligned}$$(46)
We should emphasize that we are working in a mesoscopic energy scale, i.e., the energy scale is larger compared to general relativity, such that the effect of higher order terms, e.g., \(\alpha R^{2}\) cannot be ignored. On the other hand, the energy scale is much smaller compared to the Planck scale so that the additional contributions are still subdominant, i.e., \(\alpha R<1\). This is important, in particular when one obtains the scalar field in the Einstein frame in terms of the curvature. In order to obtain a closed form expression one has to expand the potential near its minimum, and this in turn requires one to neglect higher order curvature corrections, e.g., one might neglect \(\alpha ^{2}R^{2}\) in comparison with \(\alpha R\). In a nutshell, we are working in a high curvature regime such that the effect of f(R) gravity can be felt, but it is not high enough that the Einstein–Hilbert action becomes subdominant.
Another point that requires clarification is the approximations involved in general scenarios. The conformal transformation and hence the conversion of a potential to a corresponding f(R) model is not at all straightforward. In most of the cases the relations turn out to be noninvertible, and one needs to resort to approximations. As explained earlier, on physical grounds, one can assume that the higher order terms are subleading and hence one can keep only linear order terms. While dealing with complicated potentials, most often one needs to resort to these approximations, justified by physical intuitions. However, at the Planck scale these approximations break down, since the assumption that higher orders terms are subleading cannot be trusted.
Having discussed two possible scenarios in the context of bulk physics let us now consider brane dynamics. In particular, we will be interested in one spherically symmetric and one cosmological application.
6 Jordan to Einstein frame in the brane: explicit examples
 In this example, we will start with a particular f(R) model on the brane, solve the effective field equations, and obtain a cosmological solution. Then using a conformal transformation the corresponding solution in the Einstein frame can be obtained. Solution in the Jordan frame Let us start with the f(R) model given by \(f(R)=f_{0}(RR_{0})^{\alpha }\), where \(f_{0}\) and \(R_{0}\) are constants and \(\alpha \ne 1\). The corresponding solution for the scale factor on the brane can be obtained by solving the effective field equations derived in [48, 74]. This leads to the power law solution \(a(t)\sim t^{n}\), where n is related to \(\alpha \) and the matter fields present on the brane. Converting back to Einstein frame We need to convert it back to the Einstein frame and hence obtain the corresponding solution in the scalar coupled gravity. For this choice for f(R), we obtain for the scalar fieldwhich in turn can be inverted, leading to, \(RR_{0}=\exp [a\kappa _{5}(\phi \phi _{0})]\), where \(a=\sqrt{3}/(2(\alpha 1))\) and \(\kappa _{5}\phi _{0}=(2/\sqrt{3})\ln ~(f_{0}\alpha )\). Then the potential can be determined readily, using Eq. (5), leading to$$\begin{aligned} \kappa _{5}\phi =\frac{2}{\sqrt{3}}\ln ~\left( f_{0}\alpha \right) +\frac{2}{\sqrt{3}}\left( \alpha 1\right) \ln ~\left( RR_{0}\right) , \end{aligned}$$(47)Hence the power law behavior of the f(R) theory transforms back to an exponential potential. Solution in Einstein frame The corresponding cosmological solution in the Einstein frame could be obtained by transforming the Jordan frame solution using the conformal factor. In this particular class of f(R) models, the conformal factor becomes \(\Omega =(f_{0}\alpha )^{1/3}(RR_{0})^{(\alpha 1)/3}\sim t^{2(\alpha 1)/3}\). Thus the corresponding solution in the Einstein frame is again cosmological with a new scale factor: \(\hat{a}(t)\sim t^{[3n4(\alpha 1)]/3}\). Hence the cosmological solution in the Einstein frame with an exponential potential is still a power law.$$\begin{aligned} V(\phi )=\frac{\alpha 1}{2\kappa _{5}^{2}f_{0}^{2/3}\alpha ^{5/3}}\exp \left[ \kappa _{5}a\left( \frac{5}{3}\frac{2\alpha }{3}\right) \left( \phi \phi _{0}\right) \right] ~. \end{aligned}$$(48)
 So far we have been dealing with power law f(R) theories. However, to show the applicability of our method to more general scenarios, we will consider the following f(R) model: \(R(\alpha /R)\) on the brane. This f(R) model in fourdimensional spacetime has been discussed in detail in [75], however, no such solution in the context of effective field equations exists, which by itself would be an interesting future work. However, in this work we will content ourselves by providing basic ingredients regarding this model. Furthermore, given a solution in the Jordan frame, use of a conformal transformation will lead to the corresponding solution in the Einstein frame. Solution in the Jordan frame Let us start with the above mentioned f(R) model. The corresponding solution for the scale factor on the brane can be obtained by solving the effective field equations and can be taken to be \(a(t)\sim t^{n}\), where n should be related to \(\alpha \) and the matter fields present on the brane. Converting back to Einstein frame We need to convert it back to Einstein frame and hence obtain the corresponding solution in the scalar coupled gravity. For this choice for f(R), we obtain for the scalar fieldwhich in turn can be inverted, leading to \(R=\sqrt{\alpha }(\exp [(2/\sqrt{3})\kappa _{5}\phi ]1)^{1/2}\). Then the potential can be determined readily, using Eq. (5), leading to$$\begin{aligned} \kappa _{5}\phi =\frac{2}{\sqrt{3}}\ln ~\left( 1+\frac{\alpha }{R^{2}}\right) ~, \end{aligned}$$(49)Expanding for small \(\phi \), we obtain \(V(\phi )=\sqrt{\alpha \kappa _{5}\sqrt{3}\phi /2}(1(5\kappa _{5}\phi /2\sqrt{3}))\). Hence, the negative power law behavior of the f(R) theory transforms back to a potential with \(\sqrt{\phi }\) as the leading order contribution. Solution in Einstein frame The corresponding cosmological solution in the Einstein frame could be obtained by transforming the Jordan frame solution using the conformal factor. In this particular class of f(R) model, the conformal factor becomes \(\Omega =[1+(\alpha /R^{2})]^{1/3}\sim [1+\alpha t^{4}]^{1/3}\), since \(R\sim t^{2}\). Thus for late times, \(\Omega \sim t^{4/3}\). Thus the corresponding late time solution in the Einstein frame is again cosmological with a new scale factor: \(\hat{a}(t)\sim t^{n+(8/3)}\), again a power law behavior. Hence the cosmological solution in the Einstein frame with \(\sqrt{\phi }\) potential is still a power law.$$\begin{aligned} V(\phi )=\sqrt{\alpha }\exp \left( \frac{5}{2\sqrt{3}}\kappa _{5}\phi \right) \sqrt{\exp [(2/\sqrt{3})\kappa _{5}\phi ]1}~. \end{aligned}$$(50)
 Another explicit spherically symmetric solution on the brane in the Jordan frame has been constructed in [48] by decomposing the electric part of the Weyl tensor into a dark radiation term U(r) and a dark pressure term P(r). Solution in Jordan frame The dark radiation U(r) and dark pressure P(r) act as auxiliary sources to the effective gravitational field equations [76]. A possible solution can be obtained when an “equation of state” between U(r) and P(r) is specified. For the particular choice \(2U(r)+P(r)=0\), we immediately obtain the corresponding spherically symmetric solution [48]Here \(Q_{0}\) and \(P_{0}\) correspond to constants of integration, \(\bar{\kappa }\) captures the effect of bulk spacetime, i.e., it depends on the bulk gravitational constant, and F(R), evaluated at the brane location, can be constructed from the original f(R) theory by taking an appropriate derivative. Assuming that the bulk scalar depends only on the bulk coordinates and for a f(R) theory of the form \(f(R)=R+\alpha R^{2}+\beta R^{4}\), the leading order behavior of F(R) is like an effective fourdimensional cosmological constant \(\Lambda _{4}\). Converting back to Einstein frame The corresponding scalar–tensor solution can be obtained by transforming the metric in Eq. (51) using the appropriate conformal factor: \(\Omega =(1+2\alpha R+4\beta R^{3})^{1/3}=[1+(3\kappa _{5}\phi /2)+(27\beta \kappa _{5}^{3}\phi ^{3}/64\alpha ^{3})]^{1/3}\). Since the scalar field depends on the extra coordinate only, the conformal factor evaluated on the location of the brane is just a constant. Solution in Einstein frame Since the corresponding conformal factor is just a constant it will scale the metric, which can be absorbed by rescaling of the time and radial coordinate by the conformal factor. Hence the solution in the Einstein frame would remain the same.$$\begin{aligned}&\mathrm{d}s^{2}=f(r)\mathrm{d}t^{2}+\frac{\mathrm{d}r^{2}}{f(r)}+r^{2}\mathrm{d}\Omega ^{2};\nonumber \\&f(r)=1\frac{2GM+Q_{0}}{r}\frac{3\bar{\kappa }P_{0}}{2r^{2}}+\frac{F(R)\Lambda _{4}}{3}r^{2}~. \end{aligned}$$(51)
7 Conclusions

The bulk field equations in the Einstein frame have been solved in the context of warped geometry models for two choices of the potential – quadratic and quartic. Following expectation, the warp factor behaves differently in these two scenarios, but leads to the desired exponential warping. These potential through conformal transformations are related to two f(R) models – (a) \(R +\alpha R^{2}\) and (b) \(R+\alpha R^{2}+\beta R^{4}\), respectively. From the solution in the Einstein frame we have obtained the solution in the f(R) representation as well, having a different warp factor behavior and an extra dimension dependent radion field.

Second, using the known solutions to the effective field equations in the f(R) representations we have obtained the corresponding solutions in the scalar–tensor representation. In the cosmological context, the scale factor still exhibits a power law behavior, but with a different power. The spherically symmetric solution results in mere rescaling of the coordinates.
Footnotes
Notes
Acknowledgments
Research of S.C. is funded by a SPM fellowship from CSIR, Government of India. He also thanks IACS, India for warm hospitality; a part of this work was completed there during a visit.
References
 1.N. ArkaniHamed, S. Dimopoulos, G. Dvali, The Hierarchy problem and new dimensions at a millimeter. Phys. Lett. B 429, 263–272 (1998). arXiv:hepph/9803315 ADSCrossRefGoogle Scholar
 2.I. Antoniadis, N. ArkaniHamed, S. Dimopoulos, G. Dvali, New dimensions at a millimeter to a Fermi and superstrings at a TeV. Phys. Lett. B 436, 257–263 (1998). arXiv:hepph/9804398 ADSCrossRefGoogle Scholar
 3.I. Antoniadis, A possible new dimension at a few TeV. Phys. Lett. B 246, 377–384 (1990)ADSMathSciNetCrossRefGoogle Scholar
 4.V. Rubakov, M. Shaposhnikov, Do we live inside a domain wall? Phys. Lett. B 125, 136–138 (1983)ADSCrossRefGoogle Scholar
 5.P. Horava, E. Witten, Elevendimensional supergravity on a manifold with boundary. Nucl. Phys. B 475, 94–114 (1996). arXiv:hepth/9603142 ADSMathSciNetCrossRefMATHGoogle Scholar
 6.N. Kaloper, Bent domain walls as brane worlds. Phys. Rev. D 60, 123506 (1999). arXiv:hepth/9905210 ADSMathSciNetCrossRefGoogle Scholar
 7.M. Cvetic, H.H. Soleng, Supergravity domain walls. Phys. Rep. 282, 159–223 (1997). arXiv:hepth/9604090 ADSMathSciNetCrossRefGoogle Scholar
 8.A.G. Cohen, D.B. Kaplan, Solving the hierarchy problem with noncompact extra dimensions. Phys. Lett. B 470, 52–58 (1999). arXiv:hepth/9910132 ADSMathSciNetCrossRefMATHGoogle Scholar
 9.M.D. Maia, E.M. Monte, Geometry of brane worlds. Phys. Lett. A 297, 9–19 (2002). arXiv:hepth/0110088 ADSMathSciNetCrossRefMATHGoogle Scholar
 10.I.L. Buchbinder, S.D. Odinstov, I.L. Shapiro, Effective Action in Quantum Gravity (Institute of Physics, Bristol, 1992)Google Scholar
 11.T.G. Rizzo, Higher curvature gravity in TeVscale extra dimensions. arXiv:hepph/0603242
 12.R.A. Konoplya, A. Zhidenko, (In)stability of Ddimensional black holes in GaussBonnet theory. Phys. Rev. D 77, 104004 (2008). arXiv:0802.0267 [hepth]ADSMathSciNetCrossRefGoogle Scholar
 13.R.A. Brown, Brane world cosmology with GaussBonnet and induced gravity terms. PhD thesis, Portsmouth U., ICG, 2007. arXiv:grqc/0701083
 14.T.G. Rizzo, Higher curvature effects in ADD and RS models. Pramana 69, 889–894 (2007). arXiv:hepph/0606292 ADSCrossRefGoogle Scholar
 15.R.G. Cai, N. Ohta, Black holes in pure lovelock gravities. Phys. Rev. D 74, 064001 (2006). arXiv:hepth/0604088 ADSMathSciNetCrossRefGoogle Scholar
 16.S. Chakraborty, Lanczos–Lovelock gravity from a thermodynamic perspective. JHEP 08, 029 (2015). arXiv:1505.07272 [grqc]
 17.S. Chakraborty, T. Padmanabhan, Thermodynamical interpretation of the geometrical variables associated with null surfaces. Phys. Rev. D 92(10), 104011 (2015). arXiv:1508.04060 [grqc]
 18.S. Chakraborty, N. Dadhich, BrownYork quasilocal energy in Lanczos–Lovelock gravity and black hole horizons. JHEP 12, 003 (2015). arXiv:1509.02156 [grqc]
 19.S. Chakraborty, T. Padmanabhan, Evolution of spacetime arises due to the departure from holographic equipartition in all Lanczos–Lovelock Theories of gravity. Phys. Rev. D 90(12), 124017 (2014). arXiv:1408.4679 [grqc]
 20.S. Chakraborty, T. Padmanabhan, Geometrical variables with direct thermodynamic significance in Lanczos–Lovelock gravity. Phys. Rev. D 90(8), 084021 (2014). arXiv:1408.4791 [grqc]
 21.S. Nojiri, S.D. Odintsov, S. Ogushi, Cosmological and black hole brane world universes in higher derivative gravity. Phys. Rev. D 65, 023521 (2002). arXiv:hepth/0108172 ADSMathSciNetCrossRefGoogle Scholar
 22.S. Nojiri, S.D. Odintsov, Unified cosmic history in modified gravity: from F(R) theory to Lorentz noninvariant models. Phys. Rep. 505, 59–144 (2011). arXiv:1011.0544 [grqc]
 23.T.P. Sotiriou, V. Faraoni, f(R) theories of gravity. Rev. Mod. Phys. 82, 451–497 (2010). arXiv:0805.1726 [grqc]
 24.A. De Felice, S. Tsujikawa, f(R) theories. Living Rev. Rel. 13, 3 (2010). arXiv:1002.4928 [grqc]
 25.A. Paliathanasis, \(f(R)\)gravity from killing tensors. Class. Quant. Grav. 33(7), 075012 (2016). arXiv:1512.03239 [grqc]
 26.S. Basilakos, M. Tsamparlis, A. Paliathanasis, Using the Noether symmetry approach to probe the nature of dark energy. Phys. Rev. D 83, 103512 (2011). arXiv:1104.2980 [astroph.CO]
 27.Y. Zhong, Y.X. Liu, K. Yang, Tensor perturbations of \(f(R)\)branes. Phys. Lett. B 699, 398–402 (2011). arXiv:1010.3478 [hepth]ADSCrossRefGoogle Scholar
 28.S. Capozziello, A. Troisi, PPNlimit of fourth order gravity inspired by scalartensor gravity. Phys. Rev. D 72, 044022 (2005). arXiv:astroph/0507545 ADSMathSciNetCrossRefGoogle Scholar
 29.S. Capozziello, A. Stabile, A. Troisi, The Newtonian limit of f(R) gravity. Phys. Rev. D 76, 104019 (2007). arXiv:0708.0723 [grqc]
 30.T.P. Sotiriou, The Nearly Newtonian regime in nonlinear theories of gravity. Gen. Rel. Grav. 38, 1407–1417 (2006). arXiv:grqc/0507027
 31.S. Capozziello, A. Stabile, A. Troisi, Fourthorder gravity and experimental constraints on Eddington parameters. Mod. Phys. Lett. A 21, 2291–2301 (2006). arXiv:grqc/0603071
 32.J.D. Barrow, S. Cotsakis, Inflation and the conformal structure of higher order gravity theories. Phys. Lett. B 214, 515–518 (1988)ADSMathSciNetCrossRefGoogle Scholar
 33.S. Capozziello, R. de Ritis, A.A. Marino, Some aspects of the cosmological conformal equivalence between ’Jordan frame’ and ’Einstein frame’. Class. Quant. Grav. 14, 3243–3258 (1997). arXiv:grqc/9612053
 34.S. Anand, D. Choudhury, A.A. Sen, S. SenGupta, A geometric approach to modulus stabilization. Phys. Rev. D 92(2), 026008 (2015). arXiv:1411.5120 [hepth]ADSCrossRefGoogle Scholar
 35.S. Bahamonde, S.D. Odintsov, V.K. Oikonomou, M. Wright, Correspondence of \(F(R)\) Gravity singularities in Jordan and Einstein frames. arXiv:1603.05113 [grqc]
 36.M. Parry, S. Pichler, D. Deeg, Higherderivative gravity in brane world models. JCAP 0504, 014 (2005). arXiv:hepph/0502048 ADSMathSciNetCrossRefGoogle Scholar
 37.R. Catena, M. Pietroni, L. Scarabello, Einstein and Jordan reconciled: a frameinvariant approach to scalartensor cosmology. Phys. Rev. D 76, 084039 (2007). arXiv:astroph/0604492 ADSMathSciNetCrossRefMATHGoogle Scholar
 38.T. Chiba, M. Yamaguchi, Conformalframe (In)dependence of cosmological observations in scalartensor theory. JCAP 1310, 040 (2013). arXiv:1308.1142 [grqc]
 39.S. Nojiri, S.D. Odintsov, Modified f(R) gravity consistent with realistic cosmology: From matter dominated epoch to dark energy universe. Phys. Rev. D 74, 086005 (2006). arXiv:hepth/0608008 ADSMathSciNetCrossRefGoogle Scholar
 40.S. Capozziello, S. Nojiri, S.D. Odintsov, A. Troisi, Cosmological viability of f(R)gravity as an ideal fluid and its compatibility with a matter dominated phase. Phys. Lett. B 639, 135–143 (2006). arXiv:astroph/0604431 ADSCrossRefGoogle Scholar
 41.S. Chakraborty, S. SenGupta, Higher curvature gravity at the LHC. Phys. Rev. D 90(4), 047901 (2014). arXiv:1403.3164 [grqc]
 42.S. Bhattacharya, Rotating killing horizons in generic \(F(R)\) gravity theories. arXiv:1602.04306 [grqc]
 43.S. Capozziello, M. De Laurentis, Extended theories of gravity. Phys. Rep. 509, 167–321 (2011). arXiv:1108.6266 [grqc]
 44.E. Barausse, T.P. Sotiriou, J.C. Miller, A Nogo theorem for polytropic spheres in Palatini f(R) gravity. Class. Quant. Grav. 25, 062001 (2008). arXiv:grqc/0703132
 45.S. Capozziello, M. De Laurentis, V. Faraoni, A Bird’s eye view of f(R)gravity. Open Astron. J. 3, 49 (2010). arXiv:0909.4672 [grqc]
 46.A. Aghamohammadi, K. Saaidi, M.R. Abolhasani, A. Vajdi, Spherical symmetric solution in f(R) model around charged black hole. Int. J. Theor. Phys. 49, 709 (2010). arXiv:1001.4148 [grqc]
 47.S. Capozziello, D. SaezGomez, Scalar–tensor representation of \(f(R)\) gravity and Birkhoff’s theorem. Ann. Phys. 524, 279–285 (2012). arXiv:1107.0948 [grqc]
 48.S. Chakraborty, S. SenGupta, Spherically symmetric brane spacetime with bulk \(f(\cal R\it )\) gravity. Eur. Phys. J. C 75(1), 11 (2015). arXiv:1409.4115 [grqc]
 49.N. Dadhich, R. Maartens, P. Papadopoulos, V. Rezania, Black holes on the brane. Phys. Lett. B 487, 1–6 (2000). arXiv:hepth/0003061 ADSMathSciNetCrossRefMATHGoogle Scholar
 50.T. Shiromizu, K.I. Maeda, M. Sasaki, The Einstein equation on the 3brane world. Phys. Rev. D 62, 024012 (2000). arXiv:grqc/9910076
 51.T. Harko, M. Mak, Vacuum solutions of the gravitational field equations in the brane world model. Phys. Rev. D 69, 064020 (2004). arXiv:grqc/0401049
 52.S. Chakraborty, S. SenGupta, Effective gravitational field equations on \(m\)brane embedded in ndimensional bulk of Einstein and \(f(\cal R\it )\) gravity. Eur. Phys. J. C 75(11), 538 (2015). arXiv:1504.07519 [grqc]
 53.S. Chakraborty, S. SenGupta, Spherically symmetric brane in a bulk of f(R) and GaussBonnet Gravity. arXiv:1510.01953 [grqc]
 54.I. Antoniadis, J. Rizos, K. Tamvakis, Singularityfree cosmological solutions of the superstring effective action. Nucl. Phys. B 415, 497–514 (1994). arXiv:hepth/9305025 ADSCrossRefGoogle Scholar
 55.F. Briscese, E. Elizalde, S. Nojiri, S.D. Odintsov, Phantom scalar dark energy as modified gravity: understanding the origin of the Big Rip singularity. Phys. Lett. B 646, 105–111 (2007). arXiv:hepth/0612220 ADSCrossRefGoogle Scholar
 56.M.P. Dabrowski, J. Garecki, D.B. Blaschke, Conformal transformations and conformal invariance in gravitation. Ann. Phys. 18, 13–32 (2009). arXiv:0806.2683 [grqc]
 57.S. Capozziello, P. MartinMoruno, C. Rubano, Physical nonequivalence of the Jordan and Einstein frames. Phys. Lett. B 689, 117–121 (2010). arXiv:1003.5394 [grqc]
 58.G.J. Olmo, PostNewtonian constraints on f(R) cosmologies in metric and Palatini formalism. Phys. Rev. D 72, 083505 (2005). arXiv:grqc/0505135
 59.A. Iglesias, N. Kaloper, A. Padilla, M. Park, How (Not) to Palatini. Phys. Rev. D 76, 104001 (2007). arXiv:0708.1163 [astroph]ADSCrossRefGoogle Scholar
 60.E. Barausse, T.P. Sotiriou, J.C. Miller, Curvature singularities, tidal forces and the viability of Palatini f(R) gravity. Class. Quant. Grav. 25, 105008 (2008). arXiv:0712.1141 [grqc]
 61.G.J. Olmo, Palatini approach to modified gravity: f(R) theories and beyond. Int. J. Mod. Phys. D 20, 413–462 (2011). arXiv:1101.3864 [grqc]
 62.E. Poisson, A Relativist’s Toolkit: The Mathematics of BlackHole Mechanics, 1st edn. (Cambridge University Press, Cambridge, 2007)MATHGoogle Scholar
 63.S. Nojiri, S.D. Odintsov, D. SaezGomez, Cosmological reconstruction of realistic modified F(R) gravities. Phys. Lett. B 681, 74–80 (2009). arXiv:0908.1269 [hepth]ADSMathSciNetCrossRefGoogle Scholar
 64.S. Nojiri, S.D. Odintsov, Modified gravity and its reconstruction from the universe expansion history. J. Phys. Conf. Ser. 66, 012005 (2007). arXiv:hepth/0611071 ADSCrossRefGoogle Scholar
 65.S. Nojiri, S.D. Odintsov, A. Toporensky, P. Tretyakov, Reconstruction and decelerationacceleration transitions in modified gravity. Gen. Rel. Grav. 42, 1997–2008 (2010). arXiv:0912.2488 [hepth]ADSMathSciNetCrossRefMATHGoogle Scholar
 66.S. Carloni, R. Goswami, P.K.S. Dunsby, A new approach to reconstruction methods in \(f(R)\) gravity. Class. Quant. Grav. 29, 135012 (2012). arXiv:1005.1840 [grqc]
 67.W.D. Goldberger, M.B. Wise, Modulus stabilization with bulk fields. Phys. Rev. Lett. 83, 4922–4925 (1999). arXiv:hepph/9907447 ADSCrossRefGoogle Scholar
 68.S. Chakraborty, S. Sengupta, Radion cosmology and stabilization. Eur. Phys. J. C 74(9), 3045 (2014). arXiv:1306.0805 [grqc]
 69.Y.X. Liu, Y. Zhong, Z.H. Zhao, H.T. Li, Domain wall brane in squared curvature gravity. JHEP 06, 135 (2011). arXiv:1104.3188 [hepth]ADSCrossRefMATHGoogle Scholar
 70.J.D. Barrow, S. Hervik, On the evolution of universes in quadratic theories of gravity. Phys. Rev. D 74, 124017 (2006). arXiv:grqc/0610013
 71.J.D. Barrow, S. Hervik, Simple types of anisotropic inflation. Phys. Rev. D 81, 023513 (2010). arXiv:0911.3805 [grqc]
 72.J.D. Barrow, The premature recollapse problem in closed inflationary universes. Nucl. Phys. B 296, 697–709 (1988)ADSCrossRefGoogle Scholar
 73.K.I. Maeda, Inflation as a transient attractor in R**2 cosmology. Phys. Rev. D 37, 858 (1988)ADSCrossRefGoogle Scholar
 74.Z. Haghani, H.R. Sepangi, S. Shahidi, Cosmological dynamics of brane f(R) gravity. JCAP 1202, 031 (2012). arXiv:1201.6448 [grqc]
 75.S. Nojiri, S.D. Odintsov, Modified gravity with negative and positive powers of the curvature: Unification of the inflation and of the cosmic acceleration. Phys. Rev. D 68, 123512 (2003). arXiv:hepth/0307288 ADSMathSciNetCrossRefGoogle Scholar
 76.R. Maartens, Geometry and dynamics of the brane world. in Spanish Relativity Meeting on Reference Frames and Gravitomagnetism (EREs2000) Valladolid, Spain, September 6–9, 2000. 2001. http://alice.cern.ch/format/showfull?sysnb=2237527. arXiv:grqc/0101059
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