The study of thermal and geometrical properties of black hole horizon and their intimate relationship with the dynamics of particle motion near it is one of the phenomenological interest in recent time. It has been observed that horizon may introduce chaos in a system whenever the system comes under the influence of it [1] [2] [3] [4] [5] [6] [7] [8] . In a recent work [9] , similar thing has been discussed in case of string around charged black brane. In history, chaotic dynamics in the presence of horizon has been extensively studied but the reason behind this fascinating feature of the horizon is yet to be completely understood. Similarly, why all the horizons (static or stationary) classically give the same phenomenological feature is also an essential question on this note. The investigation has not been limited in the classical scale only, and people have tried to expound the chaotic dynamics of the horizon in quantum regime as well. The phenomena of quantum chaos are mainly examined by the behaviour of the outof-time-order correlator (OTOC) of some quantum operator [10, 11] . The characteristic exponential growth of OTOC in those cases is the signature of quantum measure of chaos [10, 11] .
However, one crucial noticeable point is that, whenever we mention about chaos, there must be some instability factors associated with the system which characterises its chaotic feature. This is known as Lyapunov exponent (see [12] for a detailed discussion). One recent discovery on the upper bound of the Lyapunov exponent, predicted in Sachdev-Ye-Kitaev (SYK) model [10] , has made things very interesting. In the classical picture, it has been found in [2, 4] that for any static or stationary black holes, the radial motion of the particle grows exponentially in the near horizon region. The upper bound of * suroj176121013@iitg.ac.in † bibhas.majhi@iitg.ac.in the instability factor in this case theoretically comes out to be consistent with that of SYK model. This has been verified numerically as well 1 . In all those cases the upper bound on Lyapunov exponent is determined by the surface gravity of the black hole. Since, it is well known that the surface gravity is related to Hawking temperature [15, 16] , therefore, this upper bound is dependent on the temperature [10, 17] . In a recent paper [18] authors have demonstrated that this upper bound of the Lyapunov exponent can be verified in an experimentally realisable setup using a trapped ion technique.
In fact there are evidences about the connection between the instability of the system and its corresponding quantum thermality. The original work of M. Srednicki [19] suggested that a chaotic system naturally incorporates thermal behaviour. Recently Morita [20] in a similar note suggested that an unstable classical mode, characterised by a fixed value of Lyapunov exponent, cannot have zero temperature in the quantum scale. One of the extensively investigated unstable system in this direction is an inverse harmonic oscillator (IHO). At the classical level the IHO provides instability and people found that quantum temperature can arise from it which is determined by the instability factor. A notable feature of this analysis is that the obtained temperature is a pure quantum consequence and so in the classical limit it does vanish. All these findings indicate that the there is a close connection between the instability and the pure quantum temperature. More precisely, this instability at the classical level can be a source of a pure quantum temperature of a system.
Contemporary works of Bekenstein [21] and Hawking [15, 16] indicate that black hole horizon is a thermodynamic object. Interestingly, the horizon temperature is observer dependent quantity and more importantly is a pure quantum entity. Since this inception one of the main thrusts till date remains to look for a suitable microscopic origin of aforesaid black hole thermodynamics. There are several attempts and all of them have their own merits and demerits, and also none of them are complete.
Here we want to address one of the important issues in this area. Although the thermodynamical parameter like temperature nicely fits with the horizon, but the question remains -what is the source of this temperature? The underlying mechanism which sources such temperature still is one of the grey areas. Motivated by the earlier and recent observations in the context of connection between the classical instability and quantum thermality, we want to explore here such possibility to explain the existence of horizon temperature. We feel that it can be an important tool to explain this. In this connection we want to mention that there are some works where IHO has appeared in black hole system [2, 20, 22] . For example, Hashimoto et al [2] have shown that if one considers the analysis around the maxima of a field potential in the black hole spacetime the effective motion of a particle is that in a IHO potential. Later on Morita [20] and Hegde et al [22] independently showed that such IHO gives rise to temperature under quantization which is proportional to instability factor of the system. In a completely different context [23, 24] it has been observed that if a particle scattering phenomenon is considered in a black hole spacetime in presence of localised shock wave the effective scattering Hamiltonian comes out to be that of IHO, which also gives rise the same in the quantum regime.
The noticeable fact in all these works, mentioned above, is the possibility of the existence of instability in the form of IHO for a black hole background which provides thermality to the system. In addition, this feature is local as it exits in a very small region around a particular point -either around the maxima of the potential or around the location of the shock wave. Hence none of these analysis is directly connected with the horizon. In other words, the existing observations have not been done around the location of the horizon. Since we know that the temperature is the property of the horizon, this should arise totally from the investigations around this one way membrane. That is why our prime objective here is to find if there is any instability near the horizon, and try to understand whether that instability is associated with the this temperature.
In this regard, we mention that recently such investigation has already been attempted by us [25] . It has been observed that the near horizon Hamiltonian for the radial motion of an outgoing massless particle in the Painleve coordinates is xp type, where p is the conjugate momentum corresponding to coordinate x. It takes IHO form in a new canonical conjugate pairs of phase space variables and hence is unstable. Moreover, the consequences at the quantum level are found to be the automatic appearance of thermality as long as the Lyapunov exponent of the system remains a positive non-zero quantity. It ap-peared that the density of states (DOS) is thermal in nature with the temperature is identified as given by the Hawking expression [15] . It suggests that the instability, seen by the particle, in the classical scale around the horizon may result the horizon temperature in the quantum scale. The essential feature of this study is that the system need not to be in chaotic phase; only the unstable feature is enough to get thermality. In addition, all these are concluded with respect to the Painleve observer. Based on this fact, there are certain remaining issues which are needed to be addressed. They are as follows.
• Can the instability be addressed without going to Hamiltonian (or equivalently Lagrangian) analysis?
• Is there any other set of observers other than Painleve which also predict the similar instability and as well as thermality?
• Are all these features in general observer dependent or some are not so?
• Is it possible to construct Hamiltonian of the system just by the knowledge of the nature of instability in the near horizon regime? This will elaborate the active role of xp type Hamiltonian in this system.
• We know that the thermality of the horizon itself is an observer-dependent phenomenon. Can we predict those observers by investigating the connection between instability and thermality?
In this paper we aim to investigate the whole phenomenon in a more extensive way. In progress of addressing these issues we find that there is another set of coordinates, namely the Eddington-Finkelstein (EF) coordinates, in which the motion of the particle along null trajectory also faces the instability in the near horizon regime. Moreover, such instability is very much there for any observer when the particle is following the outgoing null path in that particular EF coordinates. It implies that the observed near horizon instability of the particle motion is an observer independent phenomena for this particular motion of the particle. Notably, again the instability factor is given by the surface gravity of the black hole. Next we find that the observer associated with the EF frame measures the radial motion as r ∼ e κt where r and t are the EF radial and time coordinates, respectively. The corresponding Hamiltonian comes out to be in xp structure. This implies that, as of now, this particular type of Hamiltonian is observer dependent -the frame (either in Painleve or in EF coordinates) which originally defines the particle's motion will see this.
Following this classical picture, we next proceed for the quantum calculation. We observe that our present EF observers are suitable to predict the automatic appearance of the thermality as a result of this aforesaid instability. We investigate this fact using different quantum approaches in order to establish our previous conjecture, stated in [25] , in a more robust way. In every cases, the temperature found out to be that given by the Hawking. Therefore now, under the present investigation, we reframe this conjecture as -
The presence of instability in the near horizon region is the mechanism for providing the temperature to the horizon as seen by a particular class of observers.
The paper is organised as follows. In section II, we first define the outgoing null path of our massless test particle in EF coordinates. We then analyse the behavior of the trajectory in section III within the classical picture in the near horizon region where the radial trajectories are found to be unstable in nature. Next using the Raychaudhuri equation [26] , in section IV we introduce a technique where we try to realise this instability in a covariant way. Section V is devoted to construct the near horizon Hamiltonian of this particle. Here first we derive it by using the unstable radial equation of motion of the particle and then verify the same using a direct approach in the context of the dispersion relation. Up to section V every calculation is done in classical scale. Now, the next section, i.e. Section VI, is dedicated to studying the quantum consequences, which is thermality of the horizon in its neighbourhood region. In subsection VI A, we start the study of thermality using tunnelling formalism [27] [28] [29] [30] [31] across the horizon. In next subsection, we investigate thermality using the detector response approach. Up next in section VII, we study the scattering of a massless particle by taking our near horizon Hamiltonian as a pure quantum mechanical scatterer. It again yields thermal nature and moreover, we are able to extract the imaginary part of the frequency of black hole quasinormal modes (QNM) [32] [33] [34] . In section VIII, we again study the thermality in the near-horizon region in a perturbative approach considering the obtained near horizon Hamiltonian as simple quantum mechanical model. In the final section (i.e. Section IX), we conclude our work. Four appendices are also included at the end of the paper. In Appendix A, we evaluate the values of the non-affinity parameter, the expansion parameter and the shear parameter for the null vector in our chosen background. These are essential for our computation of the main work. In Appendix B, we study the detector response function in (1 + 1) dimensional Schwazchild background for both the outgoing and the ingoing detector which follow the same path as our test particle. In this case the near horizon approximation is avoided. In Appendix C, we re-address the study of thermality through Gutzwiller's formula in order to strengthen our earlier analysis [25] . The last two appendices are included mainly for a side discussion.
Massless particle follows null-like trajectories and therefore the tangent to the path must be null-like.
To identify those, for simplicity, we consider a static spherically symmetric black hole (SSSBH) metric in Schwarzschild coordinates (t s , r, θ, φ) as
Usually in (1 + 3) dimensions h(r) = r 2 , but we kept this as a general function of radial coordinate for our future purpose. The above coordinate system is singular at the event horizon H, which corresponds to f (r H ) = 0. Since we shall confine our investigation to the radial path in the near horizon regime, the above singularity is not desired to exist in the choice of coordinates. Moreover, we want the particle to follow outgoing null trajectory. For this purpose, Kruskal-Szekeres (KS) coordinates (U, V, θ, φ) in null-null form will be relevant ones. Since the paths will be outgoing ones, we consider the particle propagates along the normal to U = constant surface, where
(Following the discussion in Section 2.5 of [35] , we here choose the above convention). Since the normal to null surface is tangent to it as well, the particle will propagate along this U = constant surface. For the above choice, the event horizon r = r H is located at U = 1. Here κ is the surface gravity defined by κ = f (r H )/2. The +(−) sign stands for the coordinate is defined outside (inside) the event horizon. For the present purpose, only the + sign will be considered as our test particle resides outside the horizon. In the above, u is known as the Eddington-Finkelstein (EF) outgoing null coordinate. There is also EF ingoing null coordinate v. Both of them are related to Schwarzschild coornates by the relations u = t s − r * and v = t s + r * with the tortoise coordinate r * is defined as
The KS coordinates cover the whole spacetime and therefore very much adopted to freely falling observer. In order to realize the presence of horizon by the particle and to confine it outside black hole, the null trajectories will be viewed from a different coordinate system, defined only outside the horizon. For that purpose, we adopt a new set of EF coordinates (t, r, θ, φ) where t is related to old coordinates as
where r * is taken to be valid outside the horizon. Since t s and r are timelike and spacelike in the r > r H region, these new coordinates are properly suited for this region only. The metric (1), in these, takes the following form:
Considering that the observer is in this frame, we shall calculate all our physical quantity in these coordinates. So now our next task is to calculate the normal to U = constant surface which describes the path of the massless particle. Since the observer is in new EF frame, we need to transform the normal vector in these coordinates. This will give the form of the trajectory of massless particle with respect to our desire observer. The normal vector to U = constant (say, K) surface is determined by l a = e ρ ∇ a U , where ρ is some scalar field on this. For the moment value of K can be any constant. But since we are interested in near horizon region, at the end, whenever necessary, the limit U = K → 1 will be taken to achieve our final goal. With this we find the following components of l a on any U = constant surface in (t, r, θ, φ) coordinates as
Let us now choose ρ in such a way that l t = 1. Then we obtain the contravariant components of the tangent to particle trajectory as
, 0, 0 .
Consequently the covariant components are
One can check that on the horizon H the components reduces to l a H = (1, 0, 0, 0), which has the same normalization as that of the timelike Killing vector for this spacetime. This motivated the purpose of above choice for ρ. Now the integral curves x a (µ) = (t, r, θ, φ) of l a , characterized by
where µ is parameter, fixes the particle position at a particular moment, lead to the outgoing null trajectory of our massless particle along any U = constant surface. Note that the angular components of l a vanishes and so the particle will have motion only along the radial direction. In the upcoming following section we shall study these trajectories in the near horizon regime, i.e. in the limit U → 1 (or equivalently f (r) → 0).
So far we found the path of our test particle, given by the integral curve (9) of tangent vector (7) . We are now in a position to investigate the behaviour of this curve in the vicinity of the horizon. Since it has been observed that (7) does not have any angular component, the particle will perform only the radial motion. Therefore, our local analysis will give the nature of radial coordinate of the particle.
Since the components of tangent vector l a is given by (7) and x a = (t, r, θ, φ), the time component of (9) yields
Then the radial component of (9) leads to
.
The solution of this will give us the behaviour of the particle trajectory in the radial direction. Since we are interested in the neighbourhood region of the horizon, the metric coefficient f (r) can be taken as the leading term of the Taylor series expansion of it around r = r H :
Substituting this in (11) and then keeping upto the relevant leading order (O(r − r H )), we obtain dr dt
The solution of it is
Interestingly, the above analysis indicates the presence of instability in the radial direction of the particle motion as long as the particle is very near to the horizon. Therefore we call this as local instability. For rest of the paper we shall call this as just instability without the explicit mention that it is locally applicable. But keep in mind that whenever such is stated, this is always in local sense. Before going into the discussion of the consequences of this local instability, we will show that the above can also be realised in an alternative way. We know that the expansion parameter Θ, defined in Eq. (A5) or in Eq. (A7) (see Appendix A), of the null geodesic congruence encodes the information about the behaviour of geodesics -how the distance between two neighbour paths changes. Therefore, it is instructive to investigate this parameter in this particular context. In the below we shall use the value of Θ, calculated in Appendix A, to obtain the behaviour of radial direction in the vicinity of the horizon. We shall come back to this quantity again in the next section.
Use of Eq. (A7) of Appendix A for the metric (5) yields
where prime indicates the derivative with respective to the r coordinate. Notice that, in the near horizon regime the expression (A9) for Θ implies that the expansion parameter is of the order (r − r H ). Similarly, (7) shows l r is also O(r − r H ) in this approximation. On the other hand Eq. (A4) shows thatκ = κ + O(r − r H ). Therefore, in the limit r → r H , keeping only the leading order terms in Eq. (15) we obtain
Now using the fact that l r = dr/dt, the solution of the above comes out to be r = (1/κ)e κt + C where C is an integration constant. Since, for r → r H (i.e. r * → −∞) implies t → −∞, one obtains C = r H . Therefore we have the same solution (14) . We now make a comment for the same in (1 + 1) dimension static black hole case. This is needed as later in some situation we shall consider this lower dimensional case for the simplicity of the calculation (For example, a side discussion has been made in Appendix B in this spacetime dimensions). For (1+1) dimensional case i.e. considering the (t−r) sector of metric (5), one can readily show that time and radial components of l a are given by those given in (7). Therefore, one again finds the same radial behaviour as obtained in (14) . Also as here Θ vanishes and the determinant of the metric is g = −1, the definition for expansion parameter (A7) reduces to Eq. (16) . Hence one finds (14) again and so the existence of the instability in the particle motion in the near horizon region persists in this case as well. This indicates that the present instability is completely liable for the influence of horizon in the spacetime, not a specific feature to number of spacetime dimensions.
We found that the radial motion is unstable in nature in the very near to the horizon. In this regard, it is natural to ask -what happen to the family of these null geodesics in this region? Whether these congruence of geodesics also face the similar instability due to the horizon. Moreover, in the last section we mentioned that the expansion parameter Θ can be an important quantity to illuminate our main investigation. Particularly as it measures the separation between the two nearby geodesics, it will be interesting to see how this separation changes with time. Thus we shall have a more concrete idea of instability, provided by the horizon. Therefore, the present section will be dedicated to examine the evaluation of Θ for null geodesics in the nearby region of horizon. The most promising way is to start with the Raychaudhuri's equation for null congruence [26] . Since it is in covariant form we expect that the evaluation character of Θ, obtained from this, in contrary to the earlier section, may provide a covariant description of our aforesaid instability.
Raychaudhuri equation for null geodesics is [26] 
Here we shall study this equation in the near horizon of our SSSBH spacetime (5) . All the quantities are defined with respect to the null vector (7). Let us now examine each of the terms on the right hand side of the above equation. These are all calculated in Appendix A. We found that the shear parameter σ ab = 0 (see Eq. (A12)) and since l a is hypersurface orthonormal, we must have the rotation parameter ω ab = 0 as well. Next note that in the near horizon region Θ ∼ O(r − r H ) (see Eq. (A9)), whereas as mentioned in the last section,κ = κ + O(r − r H ). The evaluation of the term R ab l a l b for metric (5) yields
This leads to the fact that R ab l a l b ∼ O(r − r H ) 2 in the near horizon approximation. Now keeping the leading order terms, i.e. O(r − r H ) terms in the right hand side of Eq. (17) one obtains
Performing the integration of the above equation we obtain the form of the expansion parameter as
This implies that in the near horizon the expansion of geodesic congruences is exponentially increasing as µ increases. It characterises the presence of instability in the geodesic motion of the particle. This analysis not only indicates the presence of local instability for the particle motion, but also provides a covariant description and realisation of this phenomenon. Since Θ is a scalar quantity, we now understand that for this particular particle motion the aforesaid instability is an observer independent feature of horizon.
Just for completeness, we now show that from Eq. (20) the explicit form of unstable nature in radial motion can be evaluated. In our EF coordinates we identified µ = t (see Eq. (10)). One can check that at the horizon (i.e. t → −∞), Θ vanishes, which implies that the above solution correctly satisfies the required boundary condition. The value of Θ for our metric (5) is given by (A9). In the near horizon regime, at the leading order, it comes out to be
where we have used h(r) = r 2 → r 2 H and h (r) = 2r → 2r H . Substitution of this in the solution (20) yields r − r H (r H /2)e κt . Thus we again we found the similar unstable nature in the radial direction. It must be mentioned that although the instability is an observer independent feature, but this particular radial character with time is related to EF observer. This is a very crucial observation in this analysis. We shall talk about more on this in our later discussion. It plays a big role in the concept of observer dependent thermality which will be introduced in the upcoming sections.
So far without using any formal prescription, like Lagrangian or Hamiltonian analysis, we have been able to show the appearance of local instability on the radial motion of a massless particle in the vicinity of horizon. This feature has been shown earlier in [25] using the Hamiltonian analysis using Painleve coordinates for the metric. It was shown that the near horizon Hamiltonian takes the form ∼ xp, where x = r − r H and p is the radial momentum. In this work we are using a new EF coordinates and found that here also similar feature is appearing in radial motion even in this new coordinates. Therefore, it would be interesting to see whether an Hamiltonian prescription can be build out in our present analysis. More importantly, we are interested to investigate the possibility of finding out the Hamiltonian of our system using the obtained radial feature in the earlier sections. If so, then whether it is again similar to xp. In this section, we shall first find the Hamiltonian from our earlier findings on radial trajectory and then verify this by deriving the same using dispersion relation for the massless particle on the background (5). This obtained structure of Hamiltonian will be very important for the later purpose of our analysis.
The near horizon radial motion is driven by Eq. (14) . Therefore use of Hamilton's equation of motionẋ = ∂H/∂p implies
where x ≡ r − r H . Solution of this is given by H = κxp+f 1 (x), where f 1 (x) is an arbitrary function of radial coordinate. This can be fixed by using the information that the corresponding Lagrangian must vanish as we are dealing with massless free particle. The Lagrangian for this solution comes out to be
So to make it vanish, we must choose f 1 (x) = 0. Thus we find that the Hamiltonian in the near horizon regime is given by
We now verify this below by a direct evaluation of Hamiltonian from the dispersion relation. This method has been adopted earlier in [25] , but for Painleve coordinates.
We again start with the static spherically symmetric metric written in EF coordinates (5) which has a timelike Killing vector χ a = (1, 0, 0, 0) and the energy of a particle moving under this background is given by E = −χ a p a = −p t , where p a is the four momentum whose components are p a = (p t , p r , 0, 0). The angular components are chosen to be zero as for our choice of path there is only radial motion (see Eq. (7)). Using the covariant form of the dispersion relation g ab p a p b = 0 for massless particle, we obtain the equation of the energy in terms of the radial component of the momentum as
It is found that the energy has two solutions:
where the positive sign for the outgoing particle and the negative sign for the ingoing one. With the near horizon approximation i.e. for f (r) given in Eq. (12), we obtain the expression for the energy of the outgoing particle (i.e. taking the +ve sign solution) as
Since, we are interested near to the horizon, taking upto the first order one obtains the expression of the Hamiltonian for the outgoing particle as (24) , with p r ≡ p. So we observed that the nature of Hamiltonian, like in Painleve coordinates, is ∼ xp even in EF coordinates. This is inherently unstable in nature having the hyperbolic points at x = 0 and p = 0 which induces the instability into the particle's motion. The solutions of the equations of motion corresponding the Hamiltonian (24) are
where λ is the affine parameter, which defines the momentum of the particle as p r = dr/dλ. It immediately shows us again that at the classical level the radial motion of the massless particle is unstable in the vicinity of the horizon which we have already shown in different approaches in the previous sections. We shall end this classical discussion with following comment. Through various approaches people have already seen that horizon may induce chaos in a system whenever the system comes under the influence of it [2, 4, 7] and notably this is common to any black hole spacetime. To follow up the real cause of this universal feature, we argued in [25] that the instability is the main cause of it. There we studied the particle motion in the Painleve coordinates. Here we showed that the same instability appears in EF coordinates. Moreover such is observer independent, provided that the particle is following a particular null path in the near horizon region. But it must be remembered that the particular radial nature of the particle trajectory is observer dependent. Since the behaviour of a system under the influence of the horizon has been studied with these trajectories, it may happen that the appearance of chaos is observer dependent phenomenon. This last statement is not conclusive at this stage; rather it is a suggestive one. We need more investigation in this direction to reach at any definite conclusion.
Till now we observed that at the classical level the horizon creates a local instability on the radial motion of a massless particle. This is completely a local phenomenon as it may not be observed when the particle motion is considered over the full spacetime. We are now curious to know whether such local phenomenon can have any observable consequence. In this section quantum aspects will be addressed. Two of the authors of this paper already showed in [25] that the instability may provide temperature to the horizon with respect to Painleve observer. Here we observed that our EF frame also perceives similar instability in the trajectory. Therefore we will again investigate if this can again explain the thermality of the horizon. Here our main objective is to study the consequences of aforesaid instability at the quantum level in various possible ways in order to verify the robustness of the aforesaid thermality. It will be found that quantum thermality of the horizon is unavoidable and thereby providing a robust evidence of our earlier claim.
Classically nothing can escape from the black hole. But the quantum probability of escaping from the barrier of horizon can be different. The previously obtained Hamiltonian can be used here to find this. It is the main quantity which is found in tunneling formalism to study the Hawking effect (For the underlying concept and details of this method see [27] [28] [29] [30] . Also see [31] for an extensive list of works on tunneling formalism) 2 . Adopting the concept of this mechanism, here we shall calculate the tunneling probability of a particle. The analysis is semi-classical in nature and calculation at the vicinity of horizon is sufficient.
We start with the standard ansatz for wave function for a particle as
where S(x) is the Hamilton-Jacobi action for the particle, defined as an integration of the momentum p of the particle with respect to the position coordinate x variable:
(Here we have considered for two dimensional phase space). The outgoing and ingoing trajectories correspond to ∂S/∂x > 0 and ∂S/∂x < 0, respectively. For our present situation both the outgoing and ingoing particles are just outside the horizon. Therefore we are interested to calculate the absorption probability of the outgoing particle while the emission probability for ingoing one. The ratio of them will give us the required tunneling probability.
The energy of the outgoing particle is given by (24) . Since H = E is the conserved quantity here, we substitute p in terms of x in (30) to find the outgoing action. Also since the absorption probability will be our main interest, the limits of the integration must be chosen x = to x = − where > 0 (i.e. from just out side the horizon to just inside). Thus the "absorption" action is given by
In performing the above integration we noticed that x = 0 is the pole of the integrand. To evaluate it the lower complex plane is being considered. Observe that since the particle starts from outside the black hole where x > 0, we have ∂S/∂x > 0 which is consistent with the definition of outgoing nature of the trajectory. On the other hand the "emission" action for the ingoing particle will be real as the limits of integration never include the horizon singularity. This can be checked trivially with the identification of energy for ingoing particle as E = −p (see Eq. (26)). So, the probability of absorption turns out to be
using the connections between the coordinates on both sides of the horizon the same has been done in [29, 30] .
where as P [Emission] = 1. Hence the tunneling probability is evaluated as
Note that the above one is thermal in nature. The temperature is identified as
This temperature exactly matches with the standard Hawking expression [15] for black hole. We just observed that the near horizon Hamiltonian (24) predicts a finite probability of escaping a particle from the horizon and thereby providing a temperature to the horizon. Since this Hamiltonian shows a local instability, we argue that such instability is responsible for the thermal behaviour of black hole. The observer is associated with the EF coordinates. But it is not vivid to us whose vacuum state is filled with particle with respect to this frame. This will be illuminated in the next subsection.
Thermality is an observer dependent phenomena [36, 37] and vacuum plays an important role in this case. The precise choice of observer and the corresponding choice of vacuum is very important in that sense. Therefore, the aim of our next approach is to identify the observer and the corresponding vacuum state corresponding this thermality. One such popular approach is investigating through the detector response of a two level atomic detector which can give us the clear idea to identify our observer and the vacuum.The choice of the observer here is the one which is following the path (14) in the near horizon regime. The vacuum is chosen to be Boulware vacua which is defined with respect to the static observer in Schwarzschild coordinates. We will find the transition rate of the atomic detector which moves along the trajectory (14) with respect to this Boulware vacuum. The calculation must be performed very near to the horizon. The particular preference of this vacuum among others like Unruh or Kruskal vacua is due to the fact that Unruh and Kruskal ones are not vacuum with respect to static frame whereas Boulware is trivial one. Therefore it is apparent that the present moving frame again find Unruh and Kruskal vacua as non-trivial one. Hence whether Boulware appears to be non-trivial with respect our present observer will be an interesting observation (a discussion of defining different vacuum states can be followed from [37] ).
Let us consider a two-level atomic detector (say a is the excited level and b is the ground state) is moving along the geodesic (14) . We consider the massless scalar field Φ under this background and its modes are denoted by u ν with frequency ν. The modes for the atomic detector we denote as ψ ω , where ω is the characteristic frequency. The interaction Hamiltonian between the atomic detector and the field is taken aŝ (35) where the operatorâ ν is the photon annihilator operator andσ ω is the atomic detector lowering operator. h.c. signifies the hermitian conjugate. Q is the coupling constant which determines the strength of the interaction and τ is the detector's clock time. This type of model was originally considered for this purpose in [38] and later has been subsequently used in [39] .
Initially, when there is no photon is detected, the detector stays in the ground state |b i.e the field is in the Boulware vacuum |0 . So the initial state of the whole system is |0, b = |0 ⊗ |b . Now after interaction the detector will go to state |a . Then the transition amplitude of the detector, using the first order perturbation theory, is given by
where |1 ν is the one particle state of Φ. In this subsection, we have chosen = 1. In this case the probability of the excitation of the atomic detector for interaction Hamiltonian (35) becomes (see [39] for detailed calculation)
Here re-expressing the detector's path (14) as t in terms of radial coordinate we obtain
where the constant, irrelevant for the present analysis, is given by (1/κ) ln(κr H ). Next taking τ = t the positive frequency mode corresponding to the detector is
The positive frequency Boulware mode for massless scalar field can be obtained by solving the Klein-Gordon (KG) equation Φ = 0 under the background of (1). Near the horizon KG equation reduces to
where in the near horizon limit r * is given by
The solutions are e −iν(ts±r * ), where the positive sign corresponds to ingoing and the negative sign refers to outgoing modes. Here, our detector is moving in the outward direction and so we will consider the ingoing Boulware mode to investigate the response of the detector. Therefore we choose
Hence expressing the integrand of Eq. (37) in terms of the radial coordinate and using (38) , we obtain the probability of transition as
where the upper limit is chosen some position r f which is situated very near to the horizon. It has to be chosen in such a way that it's value satisfies our near horizon approximation i.e. r r H − 1 << 1. In order to get some convenient look of Eq. (43) let us first make change of variable: (r/r H ) − 1 = y. Then (43) reduces to
This can be expressed in terms of lower incomplete Gamma function (See page no. 527 of ref. [40] )
But to get a better understanding, here we shall examine it numerically for different values of ω. In order to do that, first we need to make all the variables dimensionless. We choose the following substitutions in Eq. (44):
Then Eq. (44) reduces to the following form:
where P ↑ = κ 2 Q 2 r 2 H and in the above we have introduced a very small parameter to make the integration convergent. Now, we numerically integrate the above expression for different values of ω and then plot ν 2 P ↑ as a function of ν . The plot is represented in Fig. (1) . This shows that the nature of transition probability of the detector is similar to Planck distribution. So the detector will register particles in the Boulware vacuum when it moves along our local unstable path. Hence with respect to this observer the vacuum appears to be thermal. As we increase the value of ω , the peak of the curve decreases. It means that for higher values of ω , the probability of detecting particle gets lessened.
In the similar approach one can also derive the expression for probability of detecting an outgoing scalar field mode by an ingoing detector in the near horizon region. It means that the detector is moving very near to the horizon but this time its direction of motion is towards the horizon, just opposite to the previous case. In this case the EF time coordinate is represented in terms of the outgoing EF coordinates (u, r, θ, φ) as
In the similar approach we can re-express the path of the ingoing detector as t as a function of the radial coordinate. For the near horizon approximation using Eq. (B16) of the Appendix B 2 we obtain
Now, proceeding with exactly similar approach like the case of the outgoing detector one can land up to the expression of probability which turns out to be
and it basically gives the same result as in the case of the outgoing detector (FIG.(1) ). Therefore, the outgoing and the ingoing atomic detector, following the null path, detects ingoing and outgoing scalar particle, respectively in the Boulware vacuum. The Planckian nature of the plots suggest that at the quantum level the vacuum appears to be thermal. We showed this for near horizon trajectory. For completeness, we also show that our present observer when moves throughout the whole spacetime then also it will perceive thermality. This we present in Appendix B for the Schwarzschild black hole where the near horizon approximation is being avoided. To have a complete analytic analysis the calculation is performed in (1 + 1) dimensions and the temperature is identified to be the Hawking expression.
Till now we observed that the horizon provides an unstable potential to the massless particle in its neighbourhood region. Moreover, it causes the particle to feel the black hole as a thermal object. This quantum phenomenon can also be elaborated through a "scattering" model of a particle. The idea is the following. When a particle is moving very near to the horizon, it will feel the influence of the horizon through the local Hamiltonian (24) . Then the state of the particle will be influenced. The change of wave function can be evaluated by visualising (24) as the governing potential for the scattering phenomenon. In order to proceed towards the main purpose first we need to identify the initial (before scattering) and final (after scattering) energy eigenstates of the system.
Our xp Hamiltonian can be visualised as that for an inverted harmonic oscillator (IHO) in a new set of canonical variables (X, P ). The relation between the old and these new ones are x = 1 √ 2 (P − X) and p = 1 √ 2 (P + X). Then in the (X, P ) diagram, the old (x, p) variables are considered to be as ingoing and outgoing coordinates, respectively. This is shown in Fig. 2 . Here the value of x is always positive, therefore the trajectory for E > 0 in the X − P plane always remains in that quadrant where both p and x are positive definite (Fig. (2) ). Since x is identified as the ingoing coordinate, the energy eigenstate in x representation is the initial state of the system. Likewise, p representation energy state is our final state for the system. Therefore our next task is to find the eigenstates for Hamiltonian (24) in both representations.
In order to make Hamiltonian (24) hermitian, we express this asĤ
where the basic commutator is given by [x,p] = i . Then the inner product between x and p states is
To find the initial state, we represent Hamiltonian operator in position representation:
whereγ = κ. With this the initial state with energy E comes out to be
The final state is determined by expressing the Hamiltonian in momentum representation:
Eigenstate with energy E of this operator yields the final state of the system as
Now we shall find the relation between the final state and initial state. This is done as follows:
To perform the integration, we use the formula (see page no. 604 of [41] for details)
with the condition Re(b) > 0 and Re(s) > 0. To satisfy these conditions for integration (57), we take b = −i(x/ ) + and identify s = −i(E/γ) + (1/2) with > 0. At end of the integration we consider the limit → 0. This leads to
In the last step (54) has been used. So we find the relation between the final energy eigenket |E f and the initial one |E i as
The modulus square of the coefficient (C i ) in the above equation gives the probability of finding the particle in the initial state itself. Therefore, the transition probability for the particle to jump from initial (|E i ) to the final state (|E f ) is
which yields again the thermal nature with temperature is given by (34) . Now, it is well known that the scattering phenomenon in black holes can provide the information about the frequency of the QNM (see page no. 397 of [41] ). The imaginary part of the frequency is determined by the poles of the Gamma function appearing in Eq. (60). It is clear that the poles are at E n = −iγ (n + 1/2) with n = 0, 1, 2 . . . . So the imaginary part of frequency is given by ω n = −iκ(n + 1/2) which matches with the earlier finding [32] [33] [34] 41] .
Here we visualise the whole system as a following effective quantum mechanical model. We first consider a free massless particle in Minkowski spacetime whose Hamiltonian is given by H 0 = p (with the choice of unit c = 1). The near horizon Hamiltonian H κxp is treated as a small interaction of the particle with a potential of this form. So we model the actual system effectively as an interaction picture where a massless particle is interacting with the potential κxp when it is following the trajectory (14) . So we take the interaction Hamiltonian aŝ
Dirac-delta function has been introduced in order to make sure that the interaction is occurring for the particle moving along the path, given by (14) . Now if the particle is a two level quantum atom, then there is a possibility of transition from one state to another state.
Here we want to calculate the probability of transition if the atom is initially in the ground state. So the total Hamiltonian for this quantum system iŝ
whereĤ I is treated as small compared toĤ 0 . So the transition amplitude can be evaluated perturbative way. The unperturbed energy eigen basis are evaluated from H 0 =p. This will provide the initial and final basis states. These are given by ψ i (x) ∼ e iωix and ψ f (x) ∼ e iω f x , respectively (considering = 1 and the velocity of light c = 1). Introducing the transition frequency ω = ω f − ω i we write the transition amplitude at the first order perturbation as
Now, let us concentrate on
(66) The first term can be evaluated as follows:
In the similar approach the other term of Eq. (66) yields
Then, using these and then putting the values of ψ i and ψ f with their conjugates in (66) we obtain
Then, using the above expression in Eq. (65) and performing the integration one finds
Therefore the probability of transition from |i to |f turns out to be
This transition probability is thermal is nature and one identifies the temperature as (34) . Now let us give some physical aspects of this perturbation method. The main motive of this approach is to build a quantum mechanical model which mimics the near horizon characteristics. In this case we have taken the potential to be xp kind which is basically the near horizon Hamiltonian. Another important point is that we have considered a definite path for the massless particle which is basically similar to the radial trajectory of a massless particle which we have shown already in the previous sections (Sections III, IV, V). Therefore, in a physical sense this model basically mimics the behavior of the massless particle whenever it comes into the vicinity of the horizon in the quantum scale. The nonzero value of the probability whose nature is similar to the Planckian distribution tells us about the thermal behavior in the near horizon region. Therefore, it can be regarded as an effective approach to show the thermal nature in the near horizon region.
The reason why horizon is associated with temperature has always been a fascinating question towards the physics community. On the other hand, the recent observations [2, 4, 7] within the theoretical frame work predict the possibility of induction of chaotic behaviour in a system when it is under the influence of the horizon. There is a surge of discussion along this direction. Interestingly, both of these phenomenon are characterised by a common horizon quantity, namely the surface gravity. Therefore this "apparent interlink" between them may help us to uncover such properties of horizon. In our recent work [25] we predicted that the existence of local instability, created by the horizon, may be a possible reason for chaotic motion as well as horizon temperature. As a continuation, in this article we again took up this issue with great details. Here the mentioned issues in Section I have been addressed. To be concrete, we focused on a recently developed conjecture, namely the presence of local instability in the near horizon region is responsible for providing the temperature to the horizon seen by a particular set of observers. We find that this is indeed the case. For a particular set of observers, the near horizon Hamiltonian in case of an outgoing massless particle is of xp kind which is an unstable Hamiltonian, and the quantum consequences lead us to explore that thermality emerges due to this instability. This, in turn, satisfies our claim about the relationship between instability and thermality. Now, let us discuss briefly and summarize our results what we have acheived in this paper.
• In our earlier paper [25] we found that the radial motion of an outgoing massless particle in Painleve coordinate system is unstable in the near-horizon region. Proceeding one step further in this paper, our prime objective was to examine this whole oc-currence of this feature in a more extensive way. We started our calculation by choosing another set of coordinates which is EF coordinates, in which the motion of the massless particle along the null trajectory has been studied. Interestingly, we have found that in this coordinate also the instability still persists for the outgoing null path. Therefore, this result suggests that there are other sets of observers other than in Painleve coordinate system, which can also predict a similar instability.
• Moreover, in our next investigation, using the Raychaudhuri equation, we have calculated the expansion parameter of the null geodesics which are followed by our test particle in that particular EF coordinate. Here, we have found that the expansion parameter shows instability in the near horizon region. It means that in the near horizon region, instability in the particle motion is an observer independent phenomena for this particular motion of the particle.
• Following the unstable nature of the horizon, next we have constructed the Hamiltonian of the system just by the knowledge of the nature of instability in the particle motion. Interestingly, we have found that the structure of the Hamiltonian comes out to be of xp kind to the observer associated with the frame of the particle where the radial motion of the particle grows as r ∼ e κt with EF time. It suggests that although the instability is observer independent for our particle motion, the particular form of the Hamiltonian is observer dependent. The observer associated with this specific frame of the particle, either in Painleve or in EF coordinates will see this form of the near horizon Hamiltonian.
• After obtaining a clear picture of instability in the classical scale, we next targeted quantum calculations in order to see how thermality appears to our EF observers due to this unstable xp kind structure of the near horizon Hamiltonian. We started with with the tunnelling approach where we found that this near horizon Hamiltonian predicts a finite probability of escaping the particle from the horizon and thereby providing a temperature to the horizon.
• The next approach was the detector response approach in order to get a distinct idea about the relevant vacuum state. The observer or rather the detector, in this case, is following the same null trajectory in EF coordinate in the near horizon regime as we mentioned earlier. The vacuum was chosen to be that Boulware vacuum in this case. After evaluating the response function numerically, we have obtained that the transition probability of the detector of detecting a photon in the Boulware vacuum is similar to Planck distribution. It showed that the detector will see the Boulware vacuum as thermal bath.
• The other feature of the unstable potential is, it shows scattering phenomena. Therefore, our next approach was to study the scattering phenomena in presence of this unstable xp kind near horizon Hamiltonian. Indentifying the "in" and "out" states we obtained the transition probability for the particle to jump from the initial to final energy state, which yielded the thermal nature again with the desired Hawking temperature. Moreover, we gained the information about the frequency of the quasinormal modes from this scattering which matches with the earlier findings [32] [33] [34] 41] .
• In order to complete our discussion about thermality, our last approach was to consider the near horizon system as a quantum mechanical model. We built the model by considering that our near horizon xp kind Hamiltonian as a small interaction Hamiltonian, and it represents the interaction with a free massless particle. We calculated the transition probability using the first-order perturbation approximation and found out that the expression is thermal in nature. The motive of this approach was to construct a simple quantum mechanical model which mimics the near horizon characteristics. It turns out that this model definitely mimics the near horizon feature and can be regarded as an effective approach to show the thermal nature in the nearhorizon region.
Therefore, a clear recipe has been presented in this ar-ticle about the relationship between instability and the thermality in the context of the horizon. However, the specific unstable spacetime region is very small in the neighbourhood region. Moreover, it is also shown that in the quantum regime, this instability provides the automatic emergence of the temperature to the system, which is exactly equal to the Hawking expression. The way we have studied here is very interesting. Within the various known techniques the black hole system has been investigated. On this note, we feel that the results as well as the techniques, introduced here, will not only have a significant impact in the area of black hole physics, but also may uncover several unknown sides of the horizon . and l a is given by (7), Eq.(A7) yields
Next using (A4) in the above we obtain the expression of the expansion parameter Θ = h (r)f (r) h(r)(2 − f (r))
.
(A9)
Shear parameter σ ab -The shear parameter σ ab of the congruence of geodesics is defined as
where b ab is the orthogonal component of ∇ a l b projected by q ab : Substituting them in Eq. (B7) and performing the modulus square, we finally obtain the expression for the transition probability as
This thermal in nature and the temperature is identified as
which is the Hawking expression for Schwarzschild black hole.
This yielded the thermal density matrix as ρ(E) = 1 κ l 1 sinh π e − 2πE l κ cos µ l π 2 ,
with the temperature is given by (34) . See [25] for detailed calculation.
Here we shall show that the action (C4) can also be obtained in a different way. Since our Hamiltonian (24) is valid very near to the horizon, we consider a closed path which encircles the horizon x = 0 in a circular trajectory with a very small radius (say → 0) as shown in Fig. 3 . Actually the path is one which starts just outside the horizon, enters through it and comes back again. So it crosses the singular point x = 0. To avoid this a complex path has been chosen and since the relevant contribution comes from the singularity, we have chosen a circular path as shown in Fig. 3 . The choice of these types of paths motivated from the semi-classical treatment of Hawking effect in tunneling formalism, similar to what we already discussed in Section VI A. Since the formula (C1) is semi-classical in nature, we hope that such paths are relevant here as well. With this the action of the particle following the closed path is
The above closed integration can be divided into two parts:
Now, the integration I 1 i.e. when the particle is going from q A to q B is evaluated as
where in the above is chosen to be the radius of the circular path and we substituted x = e iϑ . Similarly, I 2 is evaluated to be
Finally, putting the values of I 1 and I 2 in (C7), we obtain the same expression of the action as in (C4).