Continuous THz emission from dipolaritons
We propose a scheme of continuous tunable THz emission based on dipolaritons — mixtures of strongly interacting cavity photons and direct excitons, where the latter are coupled to indirect excitons via tunnelling. We investigate the property of multistability under continuous wave (CW) pumping, and the stability of the solutions. We establish the conditions of parametric instability, giving rise to oscillations in density between the direct exciton and indirect modes under CW pumping. In this way we achieve continuous and tunable emission in the THz range in a compact single-crystal device. We show that the emission frequency can be tuned in a certain range by varying an applied electric field and pumping conditions. Finally, we demonstrate the dynamic switching between different phases in our system, allowing rapid control of THz radiation.
The development of methods for generation of continuous radiation in the terahertz range (0.3-3 THz) is currently an important physical challenge.(1) Terahertz radiation has a wide area of application, ranging from astronomy (2) and nondestructive spectroscopy (3) to security and medical imaging. Various families of devices for THz radiation generation have been theoretically proposed and experimentally tested. They include emitters based on the solid state Gunn diode (4) and the free-electron laser,(5) which allow for high output power of the radiation, while suffering from a bulky size. Microscopic devices being developed for THz generation include semiconductor structures excited with femtosecond laser pulses, where oscillations of charge density induce terahertz emission by classical dipoles.(6) However, the tunability of this source and its output power are highly restricted. Another microscopic emitter, proposed theoretically in 1971,(7) generates coherent THz radiation through repeated intersubband transitions across stacked quantum wells. These quantum cascade lasers (QCLs), realized experimentally in 1994,(8) achieve high efficiency and relatively high output power (up to 600 mW),(9) and a typical frequency of emission lying close to the upper bound of the THz range. A limiting factor of QCLs is the cryogenic temperature at which they operate. Thus, the lack of universal THz emitters with high power, efficiency, small size, which are easily-tunable and operate at relatively high temperatures stimulates the search for sources based on other operating principles.
Polaritonics — an area of physics which combines condensed matter physics and quantum optics (11); (12) — offers new possibilities. For instance, a THz emitter based on transitions between upper and lower polariton branches was considered in Refs. [(13); (14); (15)]. Another possible scheme exploits the THz range transition between 2p and 1s states of the exciton, the former being a dark state and the latter coupled to the cavity mode.(16) A recent study suggests a bosonic QCL, with multiple THz photon emission from transitions between energy levels of exciton-polaritons confined in a parabolic potential.(17)
This paper builds on the subject of a recent proposal to use a dipolariton system to generate pulses of THz radiation through oscillations in density between spatially direct and indirect excitons.(18) The dipolariton system consists of a semiconductor microcavity with a double quantum well (QW) embedded in the center (19); (20) [see sketch in Fig. 1(a)]. In a high quality cavity one can achieve strong light-matter coupling between cavity photon mode (C) and direct exciton (DX) in the left QW [Fig. 1(b)]. In the same time, tuning electron levels of the left quantum well (LQW) and right quantum well (RQW) into resonance, one can achieve the strong tunnelling coupling between a direct exciton and a spatially indirect exciton (IX) formed by an electron in the LQW and a hole in the RQW.(21); (22); (23) These strong couplings between three initial modes leads to the appearance of new eigenmodes of the system, which represent three linear superpositions of the cavity photon (C), direct exciton (DX) and indirect exciton (IX) modes. They are called the upper dipolariton (UP), the middle dipolariton (MP) and the lower dipolariton (LP).
In this paper we show that accounting for nonlinear effects arising from exciton-exciton interactions can qualitatively change the behavior of the system and allow achievement of stable continuous THz emission with tunable properties. Nonlinearities are known to give rise to bistability (24); (25); (26) when a single mode is excited with a coherent pump slightly above resonance, and multistability (27); (28) in configurations where additional states are available. Aside from the possibility of switching between different stable states,(29); (30); (31); (32) parametric instabilities can achieve periodic oscillations in particle densities.(33); (34) Here, we first investigate the property of multistability under continuous wave (CW) pumping, and the stability of the solutions. Next, we establish the conditions of parametric instability, giving rise to oscillations in density between the DX and IX modes. In this way we achieve continuous and tunable emission in the THz range, in a compact single-crystal device. We show that the emission frequency can be tuned in a certain range by an applied electric field. Finally, we demonstrate the dynamic switching between different phases in our system, allowing rapid control of THz radiation.
Ii The Model
The system is represented by two coupled QWs placed inside an optical microcavity (Fig. 1). For a coherently pumped cavity mode and resonant coupling between QWs, separated by a thin barrier, the system hosts dipolaritons — strongly mixed modes formed by cavity photon, direct exciton and indirect exciton modes. We use a generic Hamiltonian of the coupled system:
where , and are creation operators of cavity photons, direct excitons, and indirect excitons, respectively. The first three terms in Eq. (1) represent the energy of the bare cavity (), direct exciton () and indirect exciton () modes. The following two terms describe the linear coupling between modes; the first denoting the Rabi splitting between the cavity mode and the direct exciton, , and the second being the direct to indirect exciton tunneling rate . The sixth, seventh and eighth terms introduce nonlinear interactions into the system. and are the interaction matrix elements between pairs of direct and indirect excitons, respectively. The inter-species scattering of a direct exciton with an indirect exciton is described by the matrix element . The two last terms correspond to coherent pumping of the cavity mode with intensity . Under CW pumping the time dependent pumping rate can be written as , where is the energy of the pump and is its constant in time amplitude.
Equations of motion for the macroscopic order parameters defined as expectation values of annihilation operators , , can be obtained using the Heisenberg equations of motion for operators and applying the mean field approximation . Additionally we perform the change of variables . The resulting system of equations reads:
where we introduced the lifetimes of the modes ps, ns and ns. Explicit references to mode energies have been removed by defining relative energies and . The pumping term now reads . Under CW pumping we have , where , and is the relative pumping frequency.
The conditions investigated in this paper correspond to high occupation numbers of the modes for which the first order mean field approximation is applicable. To test this we also derived dynamic equations written for higher order mean field theory [see Appendix A]. Numerical solution showed no change in the results, confirming the accuracy of Eqns. (2)–(4).
In addition to dynamics we can study the steady-state properties of the system assuming a CW pump with tunable energy and intensity. It was shown for the case of a pumped mode with nonlinearity present in the system that several solutions for the occupation numbers of the modes as a function of pumping intensity are possible.(36); (37); (38); (39) Moreover, in the case of two coupled modes the system can exhibit either stable solutions or parametrically unstable solutions characterized by continuous oscillations of the mode occupation numbers.(38)
The stationary solutions can be found using the ansatz for the modes (38)
which corresponds to harmonic oscillations with the frequency of the CW pump. Inserting this into Eqns. (2)–(4) and eliminating the non-interacting photonic mode , we obtain a coupled pair of equations for the occupation numbers of the exciton modes:
where , are dressed energies, decay rates and pumping strength, expressed as
with CW solutions written in the form . To determine their stability let us explore the dynamics of a trial wave function
where is a small deviation from the stationary solution. We insert (10) into Eqns. (8)–(9), use Eqns. (6)–(7) to eliminate the pumping term , and neglect terms of second order and higher in . Finally, we get dynamic equations for deviations
where index D (I) denotes direct (indirect) exciton.
We now assume that the time dependence of the deviation can be expressed as . Inserting this ansatz in above equations yields an eigenvalue problem for , where
For each stationary state , there are four eigenvalues . The stability of the solutions can be determined from the sign of the eigenvalues. For instance, keeping in mind the form of the deviation, we see it decays in time if all imaginary parts are negative, . The solution then is stable with regards to small perturbations. If the imaginary part of one eigenvalue is non-negative, , the solution is unstable. If additionally Re, the solution is parametrically unstable.(38)
We model the system based on GaAs/AlGaAs quantum well parameters, with linear coupling parameters chosen as meV.(20) The DX-DX interaction constant can be estimated as ,(40) where the direct exciton Bohr radius and binding energy are nm and meV, respectively. Here m is the laser excitation area. The IX-IX scattering constant was taken from Ref. [(41)], with the QW separation taken as nm (4 nm tunnelling barrier). The derivation of the DX-IX interaction constant is shown in Appendix B. Relative energies of the modes were chosen as meV and meV, the latter corresponding to an applied electric field of magnitude , where kV/cm is the field strength at which the DX and IX modes are resonant.(42)
Performing numerically the stability analysis of solutions of Eqns. (6) and (7) allows us to plot a phase diagram of the system shown in Fig. 2(a). For certain values of pumping energy and pump intensity, connected to the number of indirect excitons , solutions that are not stable are possible. In these regions we can have parametric instability (PI).
Multistability curves plotted for pumping energies meV are shown in Fig. 3(a), (b) and (c,d), respectively. Stable steady state solutions are shown with narrow darker lines, and thicker lines indicate parametric instability. One can see that the DX populations are reminiscent of the conventional one mode polariton system bistability,(24); (25); (26) with regions of instability now replaced with parametric instability. This implies that population numbers do not decay to stable steady state solutions, but oscillate continuously in time. Solving numerically Eqns. (2)–(4) we show long-standing beats of IX and DX occupation numbers under CW pump (Fig. 2(b)). Note that this behavior of the system is only possible due to the presence of interactions, while without accounting for nonlinearities beats of IX-DX density quickly decay in time.(18)
The pumping energy was now fixed to meV [case (b) in Fig. 3] and the CW pump intensity was linearly turned on to s. For long turning on times ( ps) the system remains stable on the lower branches of bistability, with , and . For short turning on times ( ps) the system is excited to the parametrically unstable middle branch, resulting at large times in harmonic oscillations of the population numbers. The system can also be switched from the stable behaviour to the parametrically unstable one by applying additionally to CW pumping a short laser pulse, as shown in Fig. 4. This represents the high degree of control of the proposed system.
Furthermore, we study the behavior of the system subjected to the variation of the detuning between IX and DX modes, controlled by the applied field . The energy and intensity of the pump are kept constant. This change alters the frequency of particle numbers oscillations. The dependence of the frequency on the applied field is shown in Fig. 5, with frequency spanning from 1.5 to 2.2 THz. For the values of past oscillations become strongly anharmonic, shown in the inset in Fig. 5, which means that the system demonstrates pronounced multi-mode emission. When the electric field is decreased past the limits of Fig. 5, the time to reach harmonic oscillations increases rapidly, limiting the usefulness of this range.
We have shown that exploiting parametric instability we can achieve harmonic oscillations in the indirect exciton population under CW pumping conditions. An important property of spatially indirect excitons is a non-zero dipole moment directed in the growth direction of the microcavity. The single indirect exciton carries dipole moment equal to , where is the QW separation, while the total dipole moment of the system can be calculated as . In the case of the oscillating population numbers, this can be written:
where is the angular frequency of the IX population oscillations, and is an amplitude of the IX occupation number modulation. Thus, oscillations of the IX density induce oscillations of the total dipole moment of the system with frequency lying in the terahertz range. This results in the emission of THz radiation by the array of classical dipoles formed from dipolaritons excited over an area in the microcavity plane.
To estimate the power of emitted THz radiation, one can treat the system as a classical dipole antenna, which has a far field radiation intensity given by (43)
where is the vacuum permittivity, and is the speed of light. The radiation intensity is directionally dependent, as , where is the angle of emission with respect to the microcavity growth axis. Choosing the applied field as , the frequency of oscillations is 1.75 THz, and the maximum IX number is . The total emitted power is then nW.
It is important to note that the intensity is quadratic in the indirect exciton number. This can be explained by the phenomenon of superradiance (44); (45) — coherence of the oscillations in the quantum system causes the emitted power to increase superlinearly in the number of oscillators. The typical area over which this coherence can be realized is given by the pumping spot diameter. For example, a diameter of 60 m would give an emitted power W. The emission intensity can be further enhanced by growing additional stacks of double QWs, and by placing the system inside a supplemental THz cavity, allowing stimulated emission.
In a polariton system with GaAs/AlGaAs quantum wells, operating temperature are limited to temperatures of strong coupling observation, typically around 70 K.(35) However, one can expect that with nitride III-V compounds the operating temperature of the dipolariton emitter can be increased. All this serves to give this scheme competitive characteristics compared to other solid-state THz sources.
We have shown that nonlinear interactions in a dipolariton system give rise to multistability effects. In particular, for certain values of pumping parameters the parametric instability between IX and DX modes occurs. This results in continuous oscillations of the spatially indirect exciton occupation number under CW pumping. The frequency of these oscillations is in the THz range and can be tuned by an applied electric field. Depending on the parameters, the resulting THz radiation by the array of classical dipoles represents a continuous single-mode or multi-mode THz laser, with power output and efficiency expected to be improved over existing solid state THz emitters. We have also shown rapid switching of the THz emission controlled by an additional short laser pulse.
Acknowledgements.We would like to thank A. V. Kavokin for useful discussions. This work has been supported by FP7 IRSES project “POLATER” and AcRF Tier 1 project “Polaritonics for Novel Device Applications”. O. K. acknowledges the support from Eimskip Fund.
Appendix A Equations of motion in higher order mean field theory
In the following section we derive dynamical equations for the relevant expectation values to higher order mean field theory using the density matrix formalism. The higher order equations are then shown not to yield observable increase in accuracy.
The total Hamiltonian of the system can be written as a sum of terms describing coherent and incoherent processes, . The master equation for the density matrix reads:
where is the Lindblad superoperator.(15); (16) It accounts for dissipation processes such as non-radiative recombination of excitons, and leakage of cavity mode photons due to the non-zero transmission resonator mirrors. The explicit form of the term is
Here is the annihilation operator for the cavity mode (C), is the annihilation operator of the direct exciton (DX), and that of the indirect exciton (IX), and we omit hats of operators for simplicity. The coefficients are the decay rates of those modes ().
The equation of motion for the expectation value of an operator reads
Equations of motion can be derived using simple operator algebra, and applying the mean field approximation. For illustration purposes, let us derive the terms corresponding to the nonlinear interaction part of the Hamiltonian,
where is the scattering matrix element for the interaction between two direct excitons, is the indirect-indirect exciton interaction matrix element, and is the matrix element for scattering of a direct and an indirect exciton, calculated in Appendix B. Using the equation (18) the nonlinear part of dynamics for operator reads
Now, all number operators of the system commute with . Applying this equation to the particle numbers operators , we get
Let us now apply Eq. (20) to the correlator :
In the second to last step we used second order mean field theory to close the system of equations. Equations of motion for correlators and can similarly be calculated as
Next, we derive the dynamic equations for expectation values of the annihilation operators. Inserting the operator in Eq. (20) we get
and in the same way derive equation for ,
Note that creation or annihilation operators for the cavity mode does not appear in , and
Repeating this procedure for the other terms in the Hamiltonian leads to a closed set of equations. We perform a change of variables , . This preserves particle numbers and correlators, and removes reference to absolute mode energies. Equations can then be expressed using relative energies , , and read
For accuary comparison, numerical modelling was done in two ways. First the system of the nine coupled equation written for second order mean-field approximation were solved for certain conditions. Second the three first order equations for expectation values , and were solved for the same conditions, setting . The latter correspond to Eqns. (2)–(4) in the main text. The results of this comparison is shown in Fig. 6. The second order theory is shown only to yield observable increase in accuracy for pumping strenghts many orders of magnitude smaller than the ones treated in the paper. For this reason the first order theory was chosen to be presented in the paper, and used for obtaining the results shown.
Appendix B DX-IX interaction constant
The calculation of the matrix element of the interaction between the direct and indirect exciton follows the derivation of IX-IX interaction constant performed in Ref. [(41)]. However, the difference comes from difference of wave function of interacting particles corresponding to direct and indirect exciton, and distinct geometry of the Coulomb interaction between carriers. The matrix element can be written as the sum , where
Here is the Coulomb interaction between one direct exciton and one indirect exciton . It can be expressed as
where is a separation distance between centers of coupled QWs and we used the narrow QW approximation.
Going to the center of mass coordinates of the excitons we set , and , where , . Separating the center of mass motion and the relative motion, we can write
where corresponds to the center-of-mass momentum of the exciton. The relative motion part of wave function can be written using the variational procedure, where trial functions can be represented in several different forms, for example (46)